text:
\section{Introduction}
\subsection{Gravity and renormalizability}

It is a well know fact that general relativity is not a renormalizable theory. This means that, successful as is may be as a classical theory of gravity, it should be viewed as an effective theory that breaks down at some scale. Beyond that scale general relativity is not enough to describe the gravitational interaction or spacetime itself (depending on the perspective) and one cannot construct its quantum counterpart using conventional quantization techniques.

On the other hand, given that within this perspective general relativity is an effective theory, the Einstein-Hilbert action contains only the lowest order terms in a curvature expansion. A reasonable question is then whether a renormalizable theory can be obtained by including higher order curvature terms. Intuitively this could work, as such terms would modify the propagator of the graviton at high energies. Indeed, in 1977 Stelle showed that theories whose action includes invariants quadratic in the curvature are renormalizable \cite{stelle}. However, unfortunately this comes at a very high price, as such theories contain ghost degrees of freedom and are, therefore, not unitary. This is because by adding higher order curvature invariants in the action, one adds higher order time derivatives.

One could attempt to modify the propagator by adding higher order spatial derivatives without adding higher order time derivatives. This could presumably lead to a theory with improved ultraviolet (UV) behaviour without having to face the dreadful consequences of having higher order time derivatives. Obviously, following such a prescription requires treating spatial and time derivatives on a different footing and as such is in clash with Lorentz invariance. On the other hand, since the modified behaviour of the propagator is strictly needed in the UV, it is conceivable that Lorentz invariance could be recovered in the infrared (IR), or at least that Lorentz violations in the IR could stay below current experimental constraints.

Ho\v{r}ava's proposal \cite{Horava:2008ih,Horava:2009uw}, which has led to a spree of publications recently and is now commonly referred to as Ho\v{r}ava--Lifshitz gravity, can be seen as an attempt to put these heuristic arguments in a rigorous framework where one can test whether they are indeed valid when it comes to gravity theories. In what comes next an overview of this framework is given and its major developments to date are briefly discussed. This is far from an exhaustive account of an already very extensive literature, but hopefully it covers the most critical topics and provides a concise introduction to the subject.\footnote{See also Ref.~\cite{Weinfurtner:2010hz} for an early review of the projectable Ho\v{r}ava--Lifshitz gravity.}

\subsection{Lorentz violations as a field theory regulator}

What has been discussed above is essentially the idea of giving up Lorentz invariance in order to keep a quantum filed theory finite. Clearly, this idea goes beyond gravity theories and it has been considered in the past for other fields, see for example Ref.~\cite{anselmi}. In fact, it would be best to start from a simpler example, such as a scalar field. This section will closely follow Ref.~\cite{Visser:2009fg}.

Let us first of all point out that there is nothing wrong  with using Lorentz violations as a field theory regulator from a perspective of logical consistency, as there is no reason why a theory should necessarily exhibit Lorentz symmetry in the far UV. On the other hand, Loretz violations are severely constrained in a wide range of energies and especially in the IR. Therefore, as mentioned before, the real question is whether one can construct a field theory that exhibits Lorentz violations which in the far UV lead to renormalizability, but remain suppressed below experimental accuracy at lower energies. This is difficult to answer without concrete models and enough to motivate further investigation.

Consider the scalar field action
\begin{equation}
S_\phi=\int dt dx^d\left\{ \dot{\phi}^2-\sum_{m=1}^z a_m\phi(-\Delta)^m \phi+\sum_{n=1}^M g_n \phi^n \right\},
\end{equation}
where $\dot{}\equiv \partial/\partial t$, $\Delta=\vec{\nabla}^2$ is the spatial Laplacian and $z$ and $M$ are positive integers still to be specified.  A theory is said to be ``power counting renormalizable'' when all of its interaction terms scale like energy to some non-positive power, as in this case Feynman diagrams are expected to be convergent (or have at most a logarithmic divergence). To check whether this is true or not we first have to choose the engineering dimensions of space and time. The engineering dimensions
\begin{equation}
[dt]=[\kappa]^{-z},\qquad [dx]=[\kappa]^{-1},
\end{equation}
where $\kappa$ is a place holder symbol with dimensions of momentum, are compatible with the choice $a_z=1$ and the requirement that the action has to be dimensionless. This choice of engineering dimensions, therefore, appears to be natural in the UV. The dimensions for the scalar field then turn out to be
\begin{equation}
[\phi]=[\kappa]^{(d-z)/2}.
\end{equation}
What remain is to check the scaling for the rest of the terms. Requiring that an interaction term scales like energy to some non-positive power is equivalent to the requirement that the coupling of this interaction scales like energy to some non-negative power (since the action is dimensionless). Also, we find it convenient to work in terms of momenta instead of energies. So, our theory will be power counting renormalizable if all couplings have non-negative momentum dimensions.

It is straightforward to verify that
\begin{equation}
[a_m]=[\kappa]^{2(z-m)}, \qquad [g_n]=[\kappa]^{d+z-n(d-z)/2}.
\end{equation}
It is then obvious that $a_m$ has non negative momentum dimension for all values of $m$. It is also easy to realize that, for $z\geq d$, $g_n$ has non-negative momentum dimensions for all values of $n$. When $z<d$, $g_n$ has non-negative momentum dimension only when $n\leq[2(d+z)]/(d-z)$. One can then easily recover well know results. For example one can verify that the case $d=3$, $z=1$, $M=4$ which correspond to the usual relativistic $\phi^4$ theory in $3+1$ dimensions is a power counting renormalizable theory. The $d=z=1$ case, {\em i.e.}~the relativistic scalar in $1+1$ dimensions, is power counting renormalizable for all values of $M$. In fact this case belongs to the class $z=d$ and is directly analogous to the case $d=z=3$, which will also have the same property and is of interest to us. In fact one can argue that any theory with $z=d$ is actually finite with normal ordering, while theories with $z>d$ are finite even without normal ordering, see Ref.~\cite{Visser:2009fg} for a more detailed discussion using the ``superficial degree of divergence''.

How would things go if instead of a scalar field we wanted to consider a graviton? The main difference lies in the graviton self-interaction vertices. In the case of the scalar field considered above, momenta did not enter these interactions. For a graviton, however, one would have to deal with self-interaction vertices which contain spatial derivatives \cite{Visser:2009ys}. This introduces some further complication, but it is not enough to spoil the power-counting renormalizability, as long as $z\geq d$, {\em i.e.}~as long as the action contains operators with at least $2d$ spatial derivatives. The presence of self-interaction vertices which contain spatial derivatives is enough though to argue that, unlike the scalar theory consider above, its graviton analogue will not be finite (even with normal ordering) \cite{Visser:2009ys}.

All of the arguments presented here support the statement that field theories that contain at least $2d$ spatial derivatives in $d+1$ dimensions are power counting renormalizable. Power counting renormalizability is a strong indication that a theory is indeed renormalizable. Little progress has been made so far in the case of gravity to check renormalizability beyond power counting for such theories.

\section{From the Lifshitz scalar to non-relativistic gravity}

We proceed to consider the exact structure of a gravity theory with the characteristics mention in the previous section. We will work in $3$ dimensions from now on but generalizing to more than 3 dimensions is straightforward. The basic requirement is that the theory should have only two time derivates but at least $2d$ spatial derivatives, that is $6$ in our case. The fact that we want to have more spatial than time derivatives implies that we need to treat space and time on a different footing. This is naturally achieved by working in the Arnowitt--Deser--Misner (ADM) decomposition of spacetime
\begin{equation}
\label{admmetric}
d s^2 = - N^2 c^2 d t^2 + g_{ij}(d x^i + N^i d t) (d x^j + N^j d t).
\end{equation}
One is essentially forced to pick a preferred foliation of spacetime in order to write down the action according to the prescription laid out in the previous section. Consequently, such an action cannot be invariant under standard diffeomorphisms, as general relativity is. It can, however, be invariant under a more restricted set: foliation preserving diffeomorphisms, {\em i.e.}~space-independent time reparametrization together with time-dependent spatial diffeomorphisms
\begin{equation}
t\rightarrow \tilde{t}(t),\qquad x^i\rightarrow \tilde{x}^i(t,x^i).
\end{equation}
This is the symmetry that the action will have to respect.

It is straightforward to realize that the only covariant quantity under spatial diffeomorphisms that contains a time derivative of the spatial metric is the extrinsic curvature
\begin{equation}
K_{ij} = {1\over2N} \left\{  \dot g_{ij} - \nabla_i N_j - \nabla_j N_i \right\},
\end{equation}
which also transforms like a scalar under time reparametrization. A dot denotes differentiation with respect to the time coordinate and $\nabla_i$ is the covariant derivative associated with the spatial metric $g_{ij}$. We want the theory to be second order in time derivatives, so we need to consider terms quadratic in the extrinsic curvature. Note that there are no invariants under the symmetry mentioned above that one can construct with time derivatives of the lapse $N$ or the shift $N_j$, without including higher order time derivatives of $g_{ij}$ as well. Therefore, the most general action one can write is
\begin{equation}
S=\frac{M_{\rm pl}^2}{2}\int d^3x d t N \sqrt{g} \left\{ K^{ij} K_{ij} - \lambda K^2 -V(g_{ij},N)\right\}\, ,
\end{equation}
where $M_{\rm pl}$ is a constant which, with some foresight, we identify with the Planck mass, $g$ is the determinant of the spatial metric $g_{ij}$, and $\lambda$ is a dimensionless running coupling. $V$ generically can depend on $g_{ij}$ and $N$ and their spatial derivatives. It does not contain time derivatives, neither does it depends on the shift $N_i$, as the symmetry of the theory does not allow one to construct any suitable invariants. Finally, power counting renormalizability requires $V$ to contain terms which are at least sixth order in spatial derivatives. From now on we will restrict ourselves to theories that contain sixth order time derivatives but not higher, as these are the simplest ones which satisfy the power counting renormalizability requirement.

\section{Potential, constraints and the various versions of the theory}

Clearly there are numerous terms one could include in $V$. Different choices have lead to different versions of the theory. In this section we list these versions, briefly presenting their basic characteristics and differences. We refrain from discussing or commenting upon their viability or consistency, as this, together with their phenomenology, will be the subject of the coming section.

\subsection{Detailed balance}
Ho\v{r}ava proposed a symmetry that $V$ would have to satisfy, that would drastically reduced the number of invariants one should consider \cite{Horava:2009uw}. This symmetry is dubbed {\em detailed balance} and it is inspired by condensed matter systems. It sums up to the requirement that $V$ should be derivable from a superpotential $W$:
\begin{equation}
V=E^{ij}{\cal G}_{ijkl}E^{kl},
\end{equation}
where
\begin{equation}
E^{ij}=\frac{1}{\sqrt{g}}\frac{\delta W}{\delta g_{ij}},
\end{equation}
and
\begin{equation}
{\cal G}^{ijkl}=\frac{1}{2}\left(g^{ik}g^{jl}+g^{il}g^{jk}\right)-\lambda g^{ij}g^{kl}.
\end{equation}
The most general action one can write with $V$ satisfying the conditions above is
\begin{eqnarray}
\label{dbaction}
S_{db}=\frac{M_{\rm pl}^2}{2}\int d^3x d t N \sqrt{g} \Bigg\{ &&\!\!\!\!\!\!\!\!\!\!K^{ij} K_{ij} - \lambda K^2-\frac{\alpha^4}{M_{\rm pl}^4} C_{ij} C^{ij}+\frac{2\alpha^2\beta}{M_{\rm pl}^3}\frac{\epsilon^{ijk}}{\sqrt{g}}R_{il}\nabla_j R^l_{\phantom{a}k}-\frac{\beta^2}{M_{\rm pl}^2}R^{ij}R_{ij}\nonumber\\
&&+\frac{\beta^2}{4}\frac{1-4\lambda}{1-3 \lambda} R^2+\frac{\beta^2 \,\zeta}{1-3\lambda} R-\frac{3\beta^2 \, \zeta^2}{1-3\lambda} M^2_{\rm pl}\Bigg\}\, ,
\end{eqnarray}
where $\epsilon^{ijk}$ is the Levi-Civita symbol,
\begin{equation}
C^{ij}=\frac{\epsilon^{ikl}}{\sqrt{g}}\nabla^k\left(R^j_{\phantom{a}l}-\frac{1}{4} R \delta^j_{\phantom{a}l}\right)
\end{equation}
is the Cotton tensor, which in 3 dimensions plays the role of the Weyl tensor, and $\alpha$, $\beta$ and $\zeta$ are dimensionless couplings. Notice that there are only 3 new couplings for a total of 6 terms in $V$.

The advantages of detailed balance is that it drastically reduces the number of terms one needs to consider and that it introduces a superpotential, which might simplify quantization. On the other hand, the action above contains a term that violates parity (the fifth order operator). Additionally, as is obvious, the last term in the action, which plays the role of a cosmological constant, is restricted to have the wrong sign with respect to the one suggested by observations when $\lambda>1/3$, as pointed out in Refs.~\cite{Sotiriou:2009gy,Sotiriou:2009bx} ($\lambda$ is supposed to be close to unity, if severe Lorentz violation are not to occur).\footnote{This is a bare cosmological constant. Also, one could analytically continue the parameters as explained in Ref.~\cite{Lu:2009em}, but in this case $W$ would cease to be real.} Note that there is nothing fundamental about detailed balance, so it is safe to consider it just a simplicity assumption.

\subsection{Projectable Ho\v{r}ava--Lifshitz gravity}

Another possible simplifying restriction proposed by Ho\v{r}ava in Ref.~\cite{Horava:2009uw} is that of {\em projectability}, which is the assumption that the lapse is just a function of time. That is $N=N(t)$. Again there is no fundamental principle behind such an assumption. Ho\v{r}ava's main motivation for initially considering it was that under this assumption one has enough gauge freedom to set $N=1$, as in general relativity. This is not possible without projectability, as the symmetries of the action allow only space independent time reparametrizations, as explained earlier.

In any case, projectability drastically reduces the number of invariants $V$ can include even if detailed balance is not imposed, as all terms with spatial derivatives of $N$ would now vanish. Let us see that in more detail, following the line of Refs.~\cite{Sotiriou:2009gy,Sotiriou:2009bx}. Since $N$ is now just a function of time there is no invariant one can built out of it without using time derivatives. So $V$ will only depend on the metric and its spatial derivatives. Essentially, this means that the action should include all of the curvature invariants one can construct with $g_{ij}$ with up to six spatial derivatives.

An important simplification comes from the fact that we are in 3 spatial dimensions. Therefore, the Weyl tensor vanishes identically and the Riemann tensor can be expressed in terms of the Ricci tensor. Taking also into account Bianchi identities and ignoring surface terms, one can arrive to the conclusion that the most general action is
\begin{eqnarray}
\label{paction}
S_p&=&\frac{M_{\rm pl}^2}{2}\int d^3x d t N \sqrt{g} \Bigg\{ K^{ij} K_{ij} - \lambda K^2 -g_0 \, M_{\rm pl}^2 -g_1 R -     g_2 \,M_{\rm pl}^{-2}\,R^2 -  g_3 \, M_{\rm pl}^{-2}\, R_{ij} R^{ij} \nonumber\\
&&\qquad\qquad\qquad\qquad- g_4 \,  M_{\rm pl}^{-4}\,R^3 - g_5 \,M_{\rm pl}^{-4}\, R (R_{ij} R^{ij})- g_6 \,M_{\rm pl}^{-4}\, R^i{}_j R^j{}_k R^k{}_i \nonumber\\
&&\qquad\qquad\qquad\qquad
- g_7\,M_{\rm pl}^{-4}\, R \nabla^2 R - g_8 \,M_{\rm pl}^{-4}\, \nabla_i R_{jk} \, \nabla^i R^{jk}\Bigg\}\, ,
\end{eqnarray}
where the $g_i$ are dimensionless couplings. As long as we have not coupled matter to the theory we are free to rescale the coordinate in order to set $g_1=-1$, which is the value it has in general relativity. We then have 8 remaining couplings $g_i$ which can be used to tune the scales suppressing the various operators. The following remarks are in order:
\begin{itemize}
\item We have suppressed parity violating terms.
\item Enforcing projectability on top of detailed balance (without suppressing parity violating terms) would leave action (\ref{dbaction}) unaffected, apart from the fact that $N$ would become just a function of time. As we will see later, this is quite an important difference.
\item There are just 3 more operators in the most general projectable action than in the one with detailed balance. In this sense, detailed balance does not bring significant simplification once projectability has been assumed. For this reason, and for the drawbacks listed in the end of the previous section, we will not consider detailed balance further when assuming projectabiltity.
\item $g_0$ controls the value of the cosmological constant, which, unlike the case with detailed balance, is not restricted.
\item There are two types of Lorentz violating terms in the action. The ones in $V$, which are suppressed by some scale that can be determined by tuning the couplings $g_2$ to $g_8$, and one that comes from the kinetic part due to the fact that $\lambda$ is not necessarily equal to 1. This will be a generic feature of all versions of the theory as we will see. Clearly, the term of the second kind is far more dangerous, as it introduces Lorentz violations at low energy scales.
\item All couplings are running. One can, therefore, hope that $\lambda$ will run to 1 in the IR (or sufficiently close to it for experimental constraints to be satisfied), the rest of the operators will be heavily suppressed, and the theory will effectively reduce to general relativity. Diffeomorphism invariance and Lorentz invariance would emerge as IR (approximate) symmetries.  Whether or not this will indeed be the case requires a study of the renormalization group flow which is still pending.
\end{itemize}

\subsection{Non-projectable Ho\v{r}ava-Lifshitz gravity}

We now proceed to consider the case where neither detailed balance nor projectability are enforced. This version of the theory is called non-projectable. To avoid confusion it is worth pointing out explicitly, that the version with detailed balance can be a non-projectable version. However, in this section we wish to go beyond that. Once one abandons detailed balance, adding just some specific choice of extra terms is not really an option. Radiative corrections will generate all possible terms compatible with the symmetries of the theory and, thus, all such terms should be taken into account, much like the projectable case treated above. However, as first pointed out is Ref.~\cite{Blas:2009qj}, it is not only curvature invariants of $g_{ij}$ that one can include in $V$ in this case. One can also use the quantity
\begin{equation}
a_i=\partial_i \ln{N}\,,
\end{equation}
as contractions of it with itself or curvature terms also lead to invariants. The lowest order invariant one can construct with $a_i$ is $a^i a_i$, which comes at the same order as $R$. The action will then be of the form
\begin{eqnarray}
\label{npaction}
S_{np}&=&\frac{M_{\rm pl}^2}{2}\int d^3x d t N \sqrt{g} \Bigg\{ K^{ij} K_{ij} - \lambda K^2 +\xi R + \eta \, a_ia^i+\frac{1}{M_A^2}L_4+\frac{1}{M_B^4}L_6\Bigg\}\, ,
\end{eqnarray}
where $L_4$ and $L_6$  include all possible 4th and 6th order operators respectively that one can construct using $a_i$ and $g_{ij}$. For example, $R^2$, $(a^ia_i)^2$, $\nabla^i\nabla_i R$ and $a_ia_j R^{ij}$ are some of the terms that are included in $L_4$.\footnote{$\nabla^i\nabla_i R$ and other similar terms are not surface terms, as the $N$ in the 4-volume is space dependent.} As was the case before for $g_1$, $\xi$ can now be set to unit by a coordinate rescaling when matter is not coupled to the theory. This version of the theory was first studied in Ref.~\cite{Blas:2009qj} in order to resolve some inconsistencies in the dynamics of the non-projectable version with detailed balance, as we will discuss shortly. Although it is often referred to as an ``extension" of Ho\v{r}ava-Lifshitz gravity, this terminology is probably unfortunate: this is Ho\v{r}ava-Lifshitz gravity, {\em i.e.}~a theory constructed consistently according to the prescription described in Ho\v{r}ava's paper Ref.~\cite{Horava:2009uw} and without any restriction, such as detailed balance or projectability.

The following points are worth stressing for the non-projectable version (apart from some of the remarks already made for the projectable case and apply here as well).
\begin{itemize}
\item For simplicity the cosmological constant term has been suppressed, but it can be straightforwardly restored.
\item The scales $M_A$ and $M_B$ suppressing the higher order operators have been left arbitrary. This is in analogy with having arbitrary dimensionless couplings $g_i$ in the projectable case.
\item The number of operators in the non-projectable case is an order on magnitude larger that in the projectable case.
\item If general relativity where to be recovered in the IR, not only $\lambda$ would have to run to 1 but also $\eta$ would have to run to 0.
\end{itemize}

\subsection{Imposing further symmetries}

All of the actions presented above, which correspond to different versions of Ho\v{r}ava--Lifshitz gravity, are invariant under foliation preserving diffeomorphisms only, and not the full set of diffeomorphisms. So, even though the action is in all cases quadratic in the time derivatives of $g_{ij}$, less symmetry generically implies that more degrees of freedom will be excited. We will see explicitly in the next section that this is indeed the case, and that this can, in most cases, have undesirable consequences for the consistency and viability of the theory.

A way to do away with the extra degrees of freedom is to attempt to nontrivially extend the gauge symmetry of the theory, so that it will have as many generators per spacetime point as general relativity. It was pointed out in Ref.~\cite{Horava:2010zj} that this could be done if the action would be suitably modified to acquire an extra local $U(1)$ symmetry. More precisely in Ref.~\cite{Horava:2010zj} a step by step construction of the action led to
\begin{eqnarray}
\label{hmt}
S_{{\rm extra}\, U(1)}&=&\frac{M_{\rm pl}^2}{2}\int d^3x d t N \sqrt{g} \Big\{ K^{ij} K_{ij} - K^2 -V(g_{ij},N)\nonumber\\&&\qquad\qquad+\nu \Theta^{ij}(2 K_{ij}+\nabla_i\nabla_j \nu)-A(R-2 \Omega)/N\Big\}\, ,
\end{eqnarray}
where $\Omega$ is a constant, $A$ acts as a Lagrange multiplier, $\nu$ is an auxiliary scalar field,
\begin{equation}
\Theta^{ij}\equiv R^{ij}-\frac{1}{2} g^{ij} R+\Omega g^{ij},
\end{equation}
and $V$ should as usual contain 6th order operators and may or may not satisfy detailed balance. Note that $A$ transforms as a spatial scalar and a time vector under foliation preserving diffeomorpisms, that is
\begin{equation}
A \rightarrow A+ \dot{f} A+f \dot{A}+\xi^i\partial_i A,
\end{equation}
where $\xi^i$ is the generator. On the other hand, under a local $U(1)$ gauge transformation with generator $\varphi$, $A$ and $\nu$ transform as
\begin{eqnarray}
&&A\rightarrow A+\dot{\varphi}-N^i\nabla_i \varphi,\\
&& \nu\rightarrow \nu+\varphi.
\end{eqnarray}

See Ref.~\cite{Horava:2010zj} for more details on constructing the action, as well as for discussions on anisotropic scaling, hamiltonian formulation, linearization, etc. It was argued there that in the IR this theory has very good chances to reduce to general relativity, as the high order operators in $V$ would be suppressed (as in the the previous versions of the theory), the extra fields  in the action are just auxiliary fields, and the extra gauge symmetry leads to just the usual spin 2 graviton excitation. Moreover, it was claimed in Ref.~\cite{Horava:2010zj} that the $U(1)$ symmetry forces the coefficient $\lambda$ of $K^2$ in the action, which was a running coupling in all other versions of the theory, to be equal to 1 here, as in general relativity.

This last claim has been contested recently in Ref.~\cite{daSilva:2010bm} where it has been claimed that the extra $U(1)$ symmetry does not necessarily require $\lambda=1$ and that even for arbitrary $\lambda$ the theory propagates only a spin 2 graviton. If this is true, $\lambda$ has to remain a running coupling and flow close to $1$ in the IR for general relativistic phenomenology to be recovered according to Ref.~\cite{daSilva:2010bm}. So, at the moment, whether or not the value of $\lambda$ is fixed to 1 can be considered a matter of debate.

Another potential difficulty with this version of Horava-Lifshitz gravity has to do with matter coupling. It is straightforward to notice that the Lagrange multiplier in action (\ref{hmt}) forces $R$ to be constant. Even though this is not necessarily a problem in vacuum, it cannot be generically the case once matter is coupled to the theory. This implies that matter would have to be coupled to $A$ if the theory is to resemble general relativity at low energies, and that a suitable coupling that leads to standard phenomenology in the IR should exist. Whether this is indeed the case requires further investigation (some first steps where done in Ref.~\cite{daSilva:2010bm}).
It is also worth pointing out that action (\ref{hmt}) is susceptible to radiative corrections which could modify its form.

Since this last version of Horava-Lifshitz gravity has been proposed very recently and has not yet been studied extensively, we will not consider it further. Its cosmological aspects where recently studied in Ref.~\cite{Wang:2010wi}.

\section{Projectable version: Dynamics, consistency and low energy behaviour}
\subsection{Degrees of freedom and propagators}

We return now to the projectable version of the theory and take a closer look at its dynamics. We will consider the most general action in this version, action (\ref{paction}), {\em i.e.}~we will not enforce detailed balance but our results will include it as a subcase. The field equations for this action have been derived in detail in Ref.~\cite{Sotiriou:2009bx}. We will refrain from rewriting then here due to space limitations but we will make the following remark: since $N=N(t)$, the variation with respect to $N$ will not lead to a usual local super-Hamiltonian constraint, but to a global Hamiltonian constraint. That is, the constraint will come in the form of an integral over space. This will be a major difference with respect to the non-projectable case.

We now move straight to the linearization. If for simplicity one sets the cosmological constant term to 0, then flat space is a suitable background. After suitable gauge choices, see Ref.~\cite{Sotiriou:2009bx}, one finds that the theory propagates a spin-2 mode, the usual graviton, that now satisfies a modified dispersion relation, {\em i.e.}
\begin{equation}
\ddot{ \widetilde H}_{ij}   =  -\left[  g_1  \partial^2   + g_3  M_{\rm pl}^{-2} \partial^4   + g _8  M_{\rm pl}^{-4} \partial^6 \right] \widetilde H_{ij},
\end{equation}
where ${\widetilde H}_{ij}$ is transverse and traceless. Notice that the couplings $g_3$ and $g_8$ control the scale that the lorentz violating terms become important in the dispersion relation. Recall also that in vacuum one has the freedom to rescale the coordinates and set $g_1=-1$.

As mentioned earlier, less symmetry generically means more degrees of freedom, so since the theory is invariant under foliation preserving diffeomorphisms only, we expect to find extra excitations. This is indeed the case. There is an extra scalar degree of freedom, whose linearized dynamics are governed by the action \cite{Sotiriou:2009bx}
\begin{equation}
S^p_2=- M_{\rm pl}^2 \int  d^3x d t  \left[{1 \over c_h^2}\dot{h}^2 + h\partial^2 h+\frac{8g_2+3g_3}{M_{\rm pl}^2} h\partial^4 h -\frac{8 g_7-3 g_8}{M_{\rm pl}^4} h \partial^6 h  \right ] , \label{quad}
\end{equation}
where
\begin{equation}
c_h=\frac{1-\lambda}{3 \lambda -1},
\end{equation}
and we have set $g_1=-1$ by rescaling the coordinates, in order to guarantee the stability of the spin 2 graviton.
Given the overall minus sign in eq.~(\ref{quad}), the scalar mode is a ghost whenever $1>\lambda>1/3$. On the other hand, the scalar is classically unstable at low energies whenever $\lambda>1$ or $\lambda<1/3$ \cite{Sotiriou:2009bx}.  It is conceivable that the instability can be cut off by the higher order derivatives. If we call $M_\star$ the scale where the higher order term would take over, then the time scale of the instability would be at least $1/|c_h| M_\star$. The relation between $M_\star$ and $M_{\rm pl}$ is clearly governed by the $g_i$ coefficients in eq.~(\ref{quad}). On the other hand, if the instability is not to develop within the lifetime of the universe then $|c_h| M_\star< H_0$, where $H_0$ is the Hubble parameter.  \cite{Koyama:2009hc}. Experiments rule out modifications of Newton's law at scales above $10\mu m$ which corresponds roughly to the constraint $M_\star\gtrsim 0.1 {\rm eV}$. In turn, this implies $|1-\lambda|\lesssim 10^{-61}$. This value is clearly very low and it seems unlikely that the renormalization group flow could drive $\lambda$ to it.

In Refs.~\cite{Huang:2010rq,Wang:2010ug} linearization around a de Sitter, instead of Minkowski, background was considered. As expect, the large scale modes exhibit better behaviour. However, qualitatively the result is the same overall: $\lambda$ needs to be sufficiently close to $1$ for the instability not to be a concern. Note that requiring $\lambda$ to be very close to one is not only a potential fine tuning problem. $c_h\rightarrow 0$ as $\lambda\rightarrow 1$, which means that the low momentum phase velocity of the scalar mode will be much smaller than the speed of light for such values of $\lambda$. When a velocity of a certain mode is smaller than the speed of light in a medium, relativistic particles traveling in this medium decay via the Cherenkov process \cite{Elliott:2005va}. The fact that we observe high energy cosmic rays which need to travel a significant distance to reach us is evidence enough that such modes do not exist.

\subsection{Strong coupling}

In the previous section we only considered the linearized dynamics as described by the part of the perturbative action quadratic in $h$. However, it has been pointed out by several authors that such a perturbative treatment breaks down when $\lambda$ approaches $1$ and the scalar mode gets strongly coupled \cite{Charmousis:2009tc,Blas:2009yd,Koyama:2009hc}. To see how this comes about we give below the cubic interactions of $h$,
\begin{equation}
S^p_3=M_{\rm pl}^2\int d t d^3x \left\{h (\partial h)^2  - {2 \over c_h^4} \dot{h} \partial_i h {\partial^i \over \partial^2} \dot{h}
+  \frac{3}{2}\left[{1 \over  c_h^4}    h \left( {\partial_i \partial_j \over  \partial^2}\dot{h}\right)^2 - {(2 c_h^2+1)  \over  c_h^4}  h \dot{h}^2   \right]
  \right\} \label{cubic},
  \end{equation}
where we have neglected cubic interactions coming from higher order operators. The redefinitions $\hat{t}=|c_h| t$ and $\hat{h}=c_h^{-1/2} M_{\rm pl}\,  h$ canonically normalize the lower operator part of the quadratic action (\ref{quad}) (first two terms). After these redefinitions the cubic action reads
\begin{equation}
S^p_3=\frac{1}{|c_h|^{3/2} M_{\rm pl}}\int d \hat{t} d^3x \left\{
c_h^2\,\hat{h} (\partial \hat{h})^2  - 2 \hat{h}' \partial_i \hat{h} {\partial^i \over \partial^2}\hat{h}'
+  \frac{3}{2}\left[   \hat{h} \left( {\partial_i \partial_j \over  \partial^2}\hat{h}'\right)^2 - (2 c_h^2+1)  \hat{h}(\hat{h}')^2   \right]
  \right\} \label{cubicred},
  \end{equation}
  where now $'=\partial/\partial_{\hat{t}}$. As is obvious from the equation, there are cubic interactions which are suppressed by the scale $|c_h|^{3/2} M_{\rm pl}$ with respect to the quadratic ones. Therefore, the theory gets strongly coupled at the scale $M_{sc}=|c_h|^{3/2} M_{\rm pl}$.

Given the constraints on $\lambda$ from stability discussed previously, the strong coupling scale is phenomenologically unacceptably low, as we know that we can treat gravity perturbatively at low energies. There is also the issue of renormalizability: our arguments for power-counting renormalizability given above where based on the validity of the perturbative treatment at all energies. If there is strong coupling such arguments simply fail.

The only subtlety here is the fact that the strong coupling arguments presented so far neglect the higher order operators. However, for such operators to become important and invalidate our treatment, the strong coupling scale $M_{sc}$ would have to be higher than the scale $M_\star$ that suppresses them. Given how low $M_{sc}$ is here, this would imply that higher order operators modify the graviton dynamics at very low energies, which would be in conflict with current experimental evidence.

Since the analysis presented here is linear, one could ask whether including non-linear effects could resolve the strong coupling problem, by an analogue of the Vainstein mechanism in massive gravity \cite{vainshtein}, {\em i.e.}~a non-perturbative restoration of  the $\lambda\rightarrow 1$ limit. Recently, in Ref.~\cite{Mukohyama:2010xz} it has been claimed that this is indeed the case in spherically symmetric, static configurations. This matter deserves further attention.

\section{Non-projectable version: Dynamics, consistency and low energy behaviour}
\subsection{Degrees of freedom and propagators}

We now turn our attention to the non-projectable version of the theory. As before, we will not enforce detailed balance in order to obtain the most general results possible. However, we will discuss explicitly that case as well.
The dynamics of the spin 2 graviton are essentially the same as in the projectable version so we will not consider them again in more detail.
As in the projectable case, also here there is an extra scalar degree of freedom, and we will focus on that.

The low energy dynamics of this scalar graviton are governed by the action
\begin{equation}
S^{np}_2=- M_{\rm pl}^2 \int  d^3x d t  \left[{1 \over c_h^2}\dot{h}^2 + \frac{\eta-2}{\eta} h\partial^2 h  \right ] , \label{quadnp}
\end{equation}
which one obtains by linearizing action (\ref{npaction}) around flat space and considering only the lowest order operators. $\xi$ has been set to $1$ by a suitable coordinate rescaling. Note that $c_h$ is defined in the same way as above and, thus, it is {\em not} the low momentum phase velocity of the scalar in the non-projectable version. The square of the latter is now
\begin{equation}
c^{\prime 2}_h= c^2_h\frac{\eta-2}{\eta}.
\end{equation}
For $h$ not to be a ghost one needs $c_h^2<1$, and for $h$ to be classically stable one needs $c_{h}^{\prime 2}>0$. These conditions are satisfied whenever \cite{Blas:2009qj}
\begin{equation}
\label{npcon}
\lambda>1, \quad 0<\eta<2,
\end{equation}
or
\begin{equation}
\lambda<1/3, \quad 0<\eta<2.
\end{equation}
In the second region in the parameter space $\lambda$ is far from $1$, which is the value it has in general relativity, so we will not consider this option further.
The presence of the operator $a_i a^i$ in action (\ref{npaction}), which is a lowest order operator and contributes to the quadratic action, has drastically improved the behaviour of the scalar graviton with respect to the projectable case we examined previously. In particular, there is now a region in parameter space, described by the constraints (\ref{npcon}), for which $h$ is neither a ghost nor classically unstable \cite{Blas:2009qj}.  Even thought we have neglected higher order operators here, the behaviour we have found would remain qualitatively the same if we had included them.

Let us examine what would happen if we had enforced detailed balance. The $a^i a_i$ term in the action would not be allowed which means $\eta=0$. In this case, the coefficient of the second term in (\ref{quadnp}) blows up and the scalar field appears to freeze. Things, however, are actually more complicated, as a more detailed analysis shows \cite{Blas:2009yd}. When detailed balance is enforced, the scalar does propagate around backgrounds that are neither static nor homogeneous. In fact it appears to satisfy a first order differential equation, which constitutes a serious problem for the theory, see Ref.~\cite{Blas:2009yd} for more details. Similar conclusions have been obtained by considering the hamiltonian formulation of the theory with detailed balance \cite{Li:2009bg,Henneaux:2009zb}. In particular, the Hamiltonian constraint is not automatically preserved by time evolution (second class constraint) and the theory propagate $5/2$ half degrees of freedom, the $1/2$ corresponding to the scalar mode which satisfies a first order equation and the rest to the two graviton polarizations. Such a theory, though mathematically consistent, exhibits no time evolution and is, therefore, physically meaningless \cite{Henneaux:2009zb}.

\subsection{Strong coupling}

We just saw that, judging from the quadratic action, the non-projectable version exhibits improved scalar dynamics with respect to the projectable version, as long as detailed balance is not enforced. However, we found earlier that the scalar mode in the projectable version gets strongly coupled at unacceptably low energies, so one has to check whether this is the case in the non-projectable version as well. To do so, we write down the cubic action for the scalar \cite{Papazoglou:2009fj}:
\begin{eqnarray}
S^{np}_3&=&M_{\rm pl}^2\int d t d^3x \left\{\left(1-\frac{4(1-\eta)}{\eta^2}\right)h (\partial h)^2  - {2 \over c_h^4} \dot{h} \partial_i h {\partial^i \over \partial^2} \dot{h}     \right. \nonumber \\&& \left.~~~~~~~~~~~~~~~~~~~
+  \left(\frac{3}{2}+\frac{1}{\eta}\right)\left[{1 \over  c_h^4}    h \left( {\partial_i \partial_j \over  \partial^2}\dot{h}\right)^2 - {(2 c_h^2+1)  \over  c_h^4}  h \dot{h}^2   \right]
  \right\} \label{cubicnp}.
  \end{eqnarray}
Same as before for the projectable case, we canonically normalize the low energy quadratic action. This requires the redefinitions
\begin{equation}
t=\sqrt{\frac{\eta}{2-\eta}}\frac{\hat{t}}{|c_h|}, \qquad h=\left(\frac{\eta}{2-\eta}\right)^{1/4} \sqrt{|c_h|} \frac{\hat{h}}{M_{\rm pl}}.
\end{equation}
 Recall that we are only interested in the region of the parameter space where conditions (\ref{npcon}) are satisfied. Under the redefinitions the cubic action reads
\begin{eqnarray}
S^{np}_3&=&\frac{(2-\eta)^2}{\eta^{1/2} c^{\prime\,3/2}_h M_{\rm pl}}\int d \hat{t} d^3x \left\{
c_h^2\,\left(1-\frac{8(1-\eta)}{(2-\eta)^2}\right)\hat{h} (\partial \hat{h})^2  - 2 \hat{h}' \partial_i \hat{h} {\partial^i \over \partial^2}\hat{h}'     \right. \nonumber \\&& \left.~~~~~~~~~~~~~~~~~~~
+  \left(\frac{3}{2}+\frac{1}{\eta}\right)\left[   \hat{h} \left( {\partial_i \partial_j \over  \partial^2}\hat{h}'\right)^2 - \left(\frac{2\, \eta\, c_h^{\prime\,2}}{2-\eta}+1\right)  \hat{h}(\hat{h}')^2   \right]
  \right\} \label{cubicrednp},
  \end{eqnarray}
  where we have expressed everywhere $c_h$ in terms of the physical quantity $c'_h$. From eq.~(\ref{cubicrednp}) one directly infers that the cubic interactions are suppressed with respect to the quadratic ones by various scales $f(|\lambda-1|,\eta) M_{\rm pl}$, where $f$ is an algebraic function whose functional form depends on which terms one considers. The scalar mode would become strongly coupled at that lowest of these scales.

Both $|\lambda-1|$ and $\eta$ essentially measure deviations from Lorentz invariance which are present at arbitrarily low scales. They are to be small for the theory to avoid experimental constraints on Lorentz violations. It is rather straightforward to verify that in this case, also the scale that the theory gets strongly coupled has to be low. Instead of performing a general analysis leaving both $|\lambda-1|$ and $\eta$ small but arbitrary, we will focus on the most interesting case where $c'_h\sim 1$. This value of $c'_h$ is the one dictated by Cherenkov radiation constraints  \cite{Elliott:2005va}, see discussion above. We then have $\eta\sim|\lambda-1|$ and the strong coupling scale turns out to be $M_{sc}\sim\sqrt{\eta} M_{\rm pl}\sim\sqrt{|\lambda-1|} M_{\rm pl}$ \cite{Papazoglou:2009fj,Kimpton:2010xi}.

Since $\eta, |\lambda-1|\ll 1$ the strong coupling scale will be much lower that the Planck scale. Absence of preferred frame effects in the Solar system observations requires $\eta, |\lambda-1| \lesssim 10^{-7}$ \cite{Will:2005va}, which in turn implies that $M_{sc} \lesssim 10^{15} {\rm GeV}$. Therefore, the strong coupling scale is too high to be phenomenologically accessible from gravitational experiments. In this sense, strong coupling in non-projectable Ho\v{r}ava--Lifshitz gravity would pose no threat if the theory were to be treated as an effective theory describing gravity at low energies. On the other hand, the mere presence of strong coupling even at relatively high energies, as in this case, is enough to seriously question the power counting renormalizability of the theory \cite{Papazoglou:2009fj}. Indeed, as mentioned also earlier when considering the projectable case, renormalizability arguments were based on the assumption that a perturbative treatment can be used to arbitrarily high energies.

However, the strong coupling arguments given here neglect the role of the higher order operators in the action. As proposed in Ref.~\cite{Blas:2009ck}, a way to avoid strong coupling is to lower the scale that suppresses the higher order operators in action (\ref{npaction}) below $M_{sc}$, {\em i.e.}~impose that $M_A\sim M_B\sim M_\star$ and $M_\star<M_{sc}$. This clearly requires the introduction of a second scale, different from $M_{\rm pl}$, as well as a peculiar hierarchy of scales. Also, it does seem to involve some large dimensionless couplings as $M_{\rm pl} \gg M_{\star}$. However, as it has been argued in Refs.~\cite{Blas:2009ck,Blas:2010hb} this hierarchy of scales is technically natural.

An important consequence of this effective constraint on $M_\star$ coming from the requirement to avoid strong coupling, first pointed out in Ref.~\cite{Papazoglou:2009fj}, is that now $M_\star$ becomes bound from both above and below. Being the scale that suppresses higher order derivatives in the dispersion relations, one would expect that the larger $M_\star$ is, the better. Assuming that this same scale suppresses higher order derivatives in the matter dispersion relations (which would also be modified in a way similar to gravity), and in particular photons, places a strong lower bound on $M_\star$: The absence of a time delay in different frequency $\gamma$-rays coming from distant astrophysical sources implies that \cite{Albert:2007qk,:2009zq} $M_\star \gtrsim 10^{11}{\rm GeV}$. However, now one has also an upper bound if strong coupling is not to be a problem, and the combined set of constraints yields
\begin{equation}
10^{15} {\rm GeV} \gtrsim M_\star \gtrsim 10^{11}{\rm GeV}.
\end{equation}
There seems to be a comfortable window for $M_\star$ within which non-projectable Ho\v{r}ava--Lifshitz gravity (without detailed balance or other restrictions) avoids strong coupling without exhibiting detectable Lorentz violations, at least with current experimental accuracy. Improving the accuracy with which we can constraint both preferred frame effect in the solar system  and modifications in the dispersion relations could potentially close this window.

\subsection{Relation to Einstein-Aether theory}

Non-projectable Ho\v{r}ava--Lifshitz gravity is a theory with a preferred foliation, which will exhibit Loretz invariance violations even at low energies. This should be clear by now by inspecting the action (\ref{npaction}) and taking into account that, given the dynamics of the scalar discussed previously, one can no longer hope that $\lambda$ and $\eta$ will run to $1$ and $0$ respectively in the IR, in order for Lorentz symmetry to emerge. Note that the operators related to these couplings are of lowest order, {\em i.e.}~dimension 2.

A model theory for gravity with preferred frame effects is Einstein-aether theory \cite{Jacobson:2000xp,Jacobson:2008aj}. It is the most general theory for a unit timelike vector field coupled to gravity (but not to matter), which is second order in derivatives.
The most general action for Einstein-aether theory, up to total derivative terms and setting aside matter coupling, is
\begin{equation} \label{S}
S_{\ae} = \frac{1}{16\pi G_{\ae}}\int \sqrt{-\bar{g}}~ (-\bar{R} + L_{\ae})
~d^{4}x
\end{equation}
where $\bar{g}_{\mu\nu}$ is the 4-dimensional metric, $\bar{g}$ its determinant, $\bar{R}$ the Ricci scalar of this metric, $\bar{\nabla}_\mu$ is the covariant derivative associated with $\bar{g}_{\mu\nu}$,
\begin{equation} \label{Lae}
L_{\rm \ae} = -M^{\alpha\beta\mu\nu} \bar{\nabla}_\alpha u_\mu \bar{\nabla}_\beta u_\nu, and
\end{equation}
with $M^{\alpha\beta\mu\nu}$ defined as
\begin{eqnarray} M^{\alpha\beta\mu\nu} = c_1 \bar{g}^{\alpha\beta}\bar{g}^{\mu\nu}+c_2\bar{g}^{\alpha\mu}\bar{g}^{\beta\nu
+c_3 \bar{g}^{\alpha\nu}\bar{g}^{\beta\mu}+c_4 u^\alpha u^\beta \bar{g}_{\mu\nu}.
\end{eqnarray}
Greek indices run from $0$ to $3$, the $c_i$ are dimensionless coupling constants,
and it is assumed that $u_\mu$ is constrained to be a unit
vector, $g^{\mu\nu}u_\mu u_\nu=1$. This constraint can be explicitly imposed by the use of a lagrange multiplier, or implicitly taken into account in the variation of the action by allowing only variations that respect it.

One may wonder whether there is any relation between Einstein-aether theory and the (naive) IR limit of non-projectable Ho\v{r}ava--Lifshitz gravity (that is, a theory described by action (\ref{npaction}) without the higher order operators suppressed by $M_A$ and $M_B$). It was shown in Ref.~\cite{Jacobson:2010mx} that the latter is actually a limiting case of the former. Let us briefly go through the arguments presented there.

Starting from the action (\ref{S}), impose the restriction that the aether be hypersurface orthogonal:
\begin{equation}
\label{ho}
u_\alpha=\frac{\partial_\alpha T}{\sqrt{g^{\mu\nu}\partial_\mu T \partial_\nu T}}\,,
\end{equation}
where the unit constraint on the aether has been taken into account. $T$  will satisfy a dynamical equation. However, since it is a scalar, on shell its dynamical equation will be implied by the contracted Bianchi identity, provided that $T$ is not a constant (this can be shown for a general scalar field coupled to gravity, provided that the theory is manifestly covariant and the other fields are assumed to satisfy their field equations). However, $T$ cannot be a constant as it now defines the foliation. Thus, we are allowed to not explicitly impose the equation of motion for $T$, which in turns implies that $T$ itself can be chosen as the time coordinate. We then have
\begin{equation}
\label{ut}
u_\alpha=\delta_{\alpha T} (g^{TT})^{-1/2}=N\delta_{\alpha T}  \,.
\end{equation}
This choice selects a preferred foliation with $N$ being the lapse function of this foliation. Substituting eq.~(\ref{ut}) into the Einstein-aether theory action (\ref{S}), one gets after some algebra
\begin{equation}\label{SBPSH}
S= \frac{M_{\rm pl}^2}{2}\!\int dt d^3x \, N\sqrt{g}(K_{ij}K^{ij} - \lambda K^2
+ \xi {}^{(3)}\!R + \eta\, a_ia^i),
\end{equation}
with the following correspondence of parameters:
\begin{equation}
\label{HLpar}
\frac{1}{8\pi M_{\rm pl}^2 \,G_{\ae}}=\xi=\frac{1}{1-c_{13}}, \quad \lambda=\frac{1+c_2}{1-c_{13}},\quad \eta=\frac{c_{14}}{1-c_{13}},
\end{equation}
where we use the convention $c_{ij}=c_i+c_j$.\footnote{Note that there is a typo in the formula relating $\xi$ and $c_{13}$ in eq.~(10) of Ref.~\cite{Jacobson:2010mx}.} Clearly, action (\ref{SBPSH}) is the same as action (\ref{npaction}) without the higher order operators. Therefore, the IR limit of non-projectable Ho\v{r}ava-Lifshitz gravity is equivalent to Einstein-aether theory with a hypersurface orthogonal aether.

This equivalence provides new insights to both theories. For Einstein-aether theory it implies that there may exist a UV completion, at least for some subclasses. For non-projectable Ho\v{r}ava--Lifshitz gravity it tells us that it can be viewed as a theory were a preferred foliation is dynamically defined by a scalar field. Note that one would expect to be able to supplement Einstein-aether gravity (with a hypersurface orthogonal aether) with suitable higher order operators and construct a covariant equivalent of the full non-projectable Ho\v{r}ava--Lifshitz gravity (see Ref.~\cite{Germani:2009yt} for an early attempt).

Clearly, on account of the equivalence,  some result obtained for Einstein-aether theory go through to the IR limit of Ho\v{r}ava--Lifshitz gravity and vice versa. A characteristic example is spherically symmetric solutions, as the aether is always hypersurface orhtogonal in spherical symmetry \cite{Jacobson:2010mx,Blas:2010hb}. For example, some spherically symmetric solutions found in Ref.~\cite{Kiritsis:2009vz} for Ho\v{r}ava--Lifshitz gravity had actually already been found earlier for Einstein-aether theory \cite{Eling:2006df}. On the other hand, some important results do not go through from general Einstein-aether theory, as they hinge on the exact form of the field equations, and specifically on whether the vector is a gradient of a scalar or not. For instance, though qualitatively similar, quantitatively the PPN constraints on non-projectable Ho\v{r}ava--Lifshitz gravity cannot simply be inferred from those of Einstein-aether theory via a parameter mapping and have been obtain independently in Ref.~\cite{Blas:2010hb}.

\section{Cosmology}

As already mentioned, all versions of Ho\v{r}ava--Lifshitz gravity presented above are written in a preferred foliation, and, therefore, they do not constitute diffeomorphism invariant theories, but they are invariant under diffeomorphisms that preserve the foliation. This symmetry provides enough gauge freedom to choose
\begin{equation}
N=1,\quad N^i=0,\quad g_{ij}=a(t)^2\delta_{ij},
\end{equation}
and bring the ADM line element (\ref{admmetric}) into the usual Friedmann--Lema\^{i}tre--Robertson--Walker (FLRW) form
\begin{equation}
ds^2=-dt^2+a(t)^2 \left[\frac{dr^2}{1-k r^2}+r^2\left(d\theta^2+\sin^2 \theta d\phi^2\right)\right],
\end{equation}
under the usual cosmological assumptions of homogeneity and isotropy. Here, $a(t)$ is the scale factor and $k=+1, 0, -1$ according to whether the universe is hyperspherical, spatially flat, or hyperbolic.

Clearly, in this setting $N$ has no space dependence. Therefore, its spatial derivatives vanish. This will make the actions of the projectable and the non-projectable versions of the theory, eqs. (\ref{paction}) and (\ref{npaction}) coincide (assuming that all operators allowed by the symmetries of the theory have been consistently taken into account). The field equations do not coincide completely, as it does make a difference whether $N$ is assumed to be space-independent {\em a priori} or whether the latter is enforced as a gauge choice. The difference lies on the Hamiltonian constraint and the rest of the equations coincide. The former is global in the projectable theory and local in the non-projectable theory. So, when it comes to background cosmology, one can study both version of the theory at the same time, just taking into account the subtle issue of the Hamiltonian constraint. The dynamics of perturbation around the background will, of course, differ in the two versions, but we will not discuss perturbations here.

Cosmology in Ho\v{r}ava--Lifshitz gravity was first studied in Ref.~\cite{Kiritsis:2009sh,Sotiriou:2009bx}. A recent review of cosmology in the projectable version has recently been given in Ref.~\cite{Mukohyama:2010xz}, and as explained above, most of the results summarized there are applicable also on the non-projectable version. Here we will only give a very brief overview of background cosmology and mention in passing some key advantages of Ho\v{r}ava--Lifshitz cosmology. We refer the reader to the relevant literature for more details.

We start by presenting the field equations in an FLRW background. The supermomentum constraint is trivially satisfied, so we are left with two equations, corresponding to the first and second Friedmann equations. In the projectable version, the Hamiltonian constraint yields
\begin{equation}
\label{f1p}
\int d^3 x a^3 \left\{\frac{3\lambda-1}{2}\; {\dot a^2\over a^2}  -  {V(a)\over 6}  - {8\pi G_N \rho\over 3}\right\}=0,
\end{equation}
where $8\pi G_N\equiv M_{\rm pl}^{-2}$, $\rho\equiv-g^{-1/2} \delta S_M/\delta N$,  $S_M$ is the matter action, and
\begin{eqnarray}
V(a) = g_0\, M_{\rm pl}^2 +  {6 g_1 k  \over  a^2} + {12(3g_2+g_3)  k^2 \over M_{\rm pl}^2\,a^4}
+ {24(9 g_4 + 3g_5+ g_6) k\over M_{\rm pl}^4 \, a^6}.
\end{eqnarray}
In the non-projectable case, the Hamiltonian constraint is local and one can do away with the integral:
\begin{equation}
\label{f1np}
\frac{3\lambda-1}{2}\; {\dot a^2\over a^2}  -  {V(a)\over 6}  = {8\pi G_N \rho\over 3}.
\end{equation}
The dynamical equations yield in both cases
\begin{eqnarray}
\label{f2}
- \frac{3\lambda-1}{2} \; {\ddot a\over a}   &=&   {1\over2}\frac{3\lambda-1}{2} {\dot a^2\over a^2}
-  {1\over12 a^2} {d[V(a)\, a^3]\over d a}
+ 4\pi G_N p,
\end{eqnarray}
where $p\equiv -g^{ij}(2/N\sqrt{g})\delta S_m/\delta g^{ij}$.

Eqs.~(\ref{f1p}) and (\ref{f2}) and Eqs.~(\ref{f1np}) and (\ref{f2}) are governing background cosmology in the projectable and non-projectable versions respectively. Notice the following: In the non-projectable version one can eliminate $\dot{a}$ from eq.~(\ref{f2}) by using eq.~(\ref{f1np}) to get
\begin{eqnarray}
\label{f3np}
- \frac{3\lambda-1}{2} \; {\ddot a\over a}   &=&    -{1\over 12 a} {d[V(a) a^2] \over d a}  +{4\pi G_N\over 3} (\rho+3p),
\end{eqnarray}
Differentiating eq.~(\ref{f1np}) and subtracting from eq.~(\ref{f3np}) yields the standard conservations law
\begin{equation}
\label{mc}
\dot{\rho}+3\frac{\dot{a}}{a} (\rho+p)=0.
\end{equation}
This implies that the conservation law for matter is implied by the two (modified) Friedmann equation and, thus, imposing it separately is neither needed nor will it constrain the dynamics. This is in direct analogy to general relativity.

Things are quite different in the projectable case, given the global nature of the Hamiltonian constraint. It has been argued in Ref.~\cite{Mukohyama:2009mz} that a global constraint, such as eq.~({\ref{f1p}), is irrelevant for the local patch of the universe inside the Hubble horizon which the FLRW spacetime is supposed to approximate. Now, we could, instead of differentiating eq.~(\ref{f1p}), integrate eq.~(\ref{f2}) and ignore eq.~(\ref{f1p}) completely. One then gets
\begin{equation}
\label{f1p2}
\frac{3\lambda-1}{2}\; {\dot a^2\over a^2}  -  {V(a)\over 6}  = \frac{8\pi G_N}{3} \left(\rho+\frac{C(t)}{a^3}\right),
\end{equation}
where the functional form of $C(t)$ depends on the conservation law matter satisfies. If we eliminate $\dot{a}$ from eq.~(\ref{f2}) by using eq.~(\ref{f1p2}) we get
\begin{eqnarray}
\label{f3p}
- \frac{3\lambda-1}{2} \; {\ddot a\over a}   &=&    -{1\over 12 a} {d[V(a) a^2] \over d a}  +{4\pi G_N\over 3} \left(\rho+\frac{C(t)}{a^3}+3p\right).
\end{eqnarray}
Now, the conservation law for matter has become essential to the dynamics via the presence of $C(t)$. The terms containing $C(t)$ are the only difference between eqs.~(\ref{f1p2}) and (\ref{f3p}), which describe the background dynamics in the projectable case, and eqs.~(\ref{f1np}) and (\ref{f3np}), which describe the background dynamics in the non-projectable case.\footnote{This analysis, leading to the same conclusions, can be performed also before imposing the FLRW ansatz and the relevant symmetries, see Ref.~\cite{Mukohyama:2009mz}.}

Now that we have the field equations in a suitable form for both cases we can proceed to discuss phenomenology. Suppose that matter is coupled in such a way that at low energies $\rho$ and $p$ have the usual meaning and satisfy the standard conservation law of eq.~(\ref{mc}) (this is a strong assumption, see also next section). For the projectable case this implies that $C(t)=C_0$, where $C_0$ is a constant. The $C_0$ related terms then play the role of a dark matter component \cite{Mukohyama:2009mz}. This dark matter component will not exist in the non-projectable case.

Setting aside the role of matter, the main difference between Ho\v{r}ava--Lifshitz gravity and general relativity in background cosmology is the presence of the last two terms in $V$. The first one is a what is known as a dark radiation component, as it scales as $a^{-4}$. The second one is what one could call a stiff matter component, as it scales as $a^{-6}$. Note that the sign of the $a^{-6}$ term depends on the sign of the $k$ and that both terms vanish when $k=0$. An interesting feature related to the presence of these terms is that they can lead to solutions with classical cosmological bounces, provided that their coefficients will have appropriate signs \cite{Sotiriou:2009bx,Kiritsis:2009sh}.

One last important property of Ho\v{r}ava--Lifshitz cosmology which is worth mentioning before closing this brief overview, is that is seems to lead to a scale invariant spectrum of cosmological perturbations, without the need for inflation \cite{Mukohyama:2009gg}. This property is related to the anisotropic scaling between space and time that the theory exhibits at high energies, which is also the key ingredient for power counting renormalizability.

\section{Conclusions and future perspectives}

A brief overview of Ho\v{r}ava--Lifshitz gravity was given and special attention was payed in distinguishing between the various different versions. After that, the main focus of this review was on the dynamics and the viability of these versions. In summary, the projectable version has serious viability problems because it harbours a scalar mode which is either classically or quantum mechanically unstable (depending on the position in parameter space) and also exhibits strong coupling at low energies (whether or not the Vainshtein mechanism can alleviate the latter is not clear yet, see also above). The non-projectable version, on the other hand, seems to have similar (and other) problems if detailed balance is imposed. These problems, however, appear to be overcome once detailed balanced is abandoned and all operators allowed by the symmetries of the theory are consistently taken into account. Alleviating the strong coupling problem requires the introduction of a second scale in the theory which should be parametrically smaller than the Planck scale, and should be the scale at which higher order operators become important. The dynamics of the version exhibiting the extra $U(1)$ symmetry, which was very recently proposed, have not been adequately studied yet.

There are many open issues regarding Ho\v{r}ava--Lifshitz gravity, the two most important ones being the following. Firstly, in the introduction we argued that the theory is power-counting renormalizable, as is usually done in the literature. Even though this is a strong indication for UV completeness,  renormalizability beyond power counting has not been explicitly shown. Additionally, the renormalization group flow of the various couplings has not been studied, which implies that, for the time being, one does not really know if the theory approaches GR in the IR ($\lambda\to 1$, $\eta\to 0$) or not. Secondly, the role of matter and its coupling to gravity have not been fully clarified yet. The matter action will have to include higher order spatial derivatives, which implies that there will be modifications in the dispersion relations of matter fields that can lead to serious constraints. Additionally, couplings between the matter and the scalar graviton could lead to violations of the equivalence principle. Even if initially omitted, such couplings would be generically generated by radiative corrections, unless some symmetry prohibits them ({\em e.g.}~supersymmetry \cite{GrootNibbelink:2004za}). Hopefully, future work will shed some light onto these issues.

\ack

I am indebted to my collaborators Antonios Papazoglou, Matt Visser and Silke Weinfurtner  for their hard work and invaluable input in our joint projects on Ho\v{r}ava--Lifshitz gravity. I would also like to thank Ted Jacobson, Ian Kimpton and Tony Padilla for numerous stimulating and enlightening discussions. Many thanks also to the participants of the Peyresq 2010 meeting. This work was supported by a Marie Curie Incoming International Fellowship.

\section*{References}

redpajama_set_name: RedPajamaArXiv


Calendars are how people visualize their day.
When dealing with multiple schedules, for potentially multiple employees, each time of which may or may not have bookings for them already, calendars can be daunting though. That's what we have worked hard to avoid with our latest new feature announcement: FlexBooker now has calendars to help you manage your appointments and bookings.
We think our implementation is, well, beautifully simple (and simply beautiful). The calendar can be seen by managers in its entirety (either day or week) and can also be filtered by employee. It is also handily color-coded, so you can see at a glance who has appointments, and where the bookings lie.
Of course, it's still amazingly fast. We've also combined our schedule setup area into the same page as the calendar, so you can perform all your day-to-day activities from one page.
We think you're going to love the new look. The search functionality is still there, so you can easily search your entire schedule, even by one of your customer's names, to quickly see his history and upcoming bookings (still on the same page!).
As always, now back to work to continually improve FlexBooker for you.
Responsive and Helpful Support. By both email and phone.
Everyone claims to have it, but we really strive to give our customers the best customer support possible. We don't like to work in days, or hours, but instead in minutes to respond to a customer question.
The same drive runs throughout the company, from support to sales. Everywhere we interact with customers, we want the customer to come away with a true "wow" feeling.
We're excited to announce a brand new FlexBooker feature: Wait Lists.
What are FlexBooker Wait Lists?
FlexBooker Wait Lists are a powerful way to help keep your appointmets full, allowing your customers to place themselves onto a wait list for any full appointment.
A wait list is not a confirmation or a booking, but is instead a way to automatically get a spot if space opens up for them.
In the event of space opening up for a wait list party, FlexBooker automatically sends the customer a text message, allowing them to turn their wait list request into a confirmed booking.
What can cause space to open up for my customers who are on a wait list for a particular time?
Is space held for wait list parties?
Yes, for a period of exactly 10 minutes from the moment space opens up (at which time the wait list customer will receive a text message asking if they would like the booking) space will be held for that customer.
If they don't respond within that 10 minute time frame, then the space goes back to the widget to allow your other customers to book.
What happens if multiple customers are waiting for space at the same appointment?
If there are multiple customers waiting for space at the same booking, then they will be offered space as (and if) it opens up in the order that they joined the wait list. First come, first offered.
If a party above them in the wait list doesn't respond to the SMS within 10 minutes, or rejects the SMS offer, then the next party in the wait list will be text messaged.
This is another great evolution to the FlexBooker online scheduling offering, and we're so happy to be able to now offer this often-requested feature to our customers. Not a customer yet? Check out what all the fuss is about.
Now, we thought, what's true of gyms should apply to other types of business.
Over the last several months we've been working closely with folks in several other businesses to adapt FlexBooker to work for them too.
Rather than trying to come out with a one-size-fits-all approach (as some of the other guys do), where each user has to spend hours fiddling through endless settings, we do the work for you. By choosing one of the business types (and we have quite a few now, and are working to grow the list as we learn more and more) we will automatically configure your account for you. The feedback so far as been great, and we truly appreciate how open and helpful so many of our customers have been. Give us a try, for 60 days(!) completely free of charge (no credit card required) by signing up for an instant free trial.
We feel strongly that we offer the best solution out there for businesses looking to accept bookings online. But we're not resting, we have lots of awesome features and improvements in the pipeline. More to come on those in upcoming posts.


As a Google Certified Teacher, Google is completely integrated into all aspects of my personal and professional life. However...I have found Weebly sites to be quick and easy to throw together.
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text:
McClellan didn't witness the accident. She said her 37-year-old brother, Josiah Bigelow, noted that he took the spark plug out of the machine before the fire occurred.
The father of five acted quickly to extinguish the flames that enveloped his eldest child, ripping off his own shirt before taking the boy to the ground to smother the flames, McClellan said. He then grabbed a fire extinguisher to put out the lawnmower fire before calling 911 to alert authorities that he was rushing his son to the hospital.
The longtime former ambulance driver for the Idaho Falls Fire Department, who also suffered severe hand and arm burns in the accident, drove his son 15 miles to Eastern Idaho Regional Medical Center in Idaho Falls. Soon after they arrived there, the boy and his father were flown by air ambulance to the burn center in Utah, McClellan aid.
Braxton has first-, second- and third-degree burns on about 45 percent of his body, including his right arm, his torso and upper thighs. He has some blistering on his face.
"I think my brother yelled at him to cover his face, and it saved his lungs and airway," McClellan said. "But it got his arms."
She said the boy is being kept in a medically-induced coma. Doctors have already removed some of the dead skin, and they'll remove more next week.
Family members are caring for Josiah and Angie Bigelow's four children while they are at Braxton's side at the burn center. The boy is expected to be in treatment there for another 45 days and in recovery for a full year.
Two teachers at Rigby Middle School are collecting cards, posters and gifts for Braxton, whose 13th birthday is Monday. The goal is to have his room decorated with get well signs and birthday wishes for when he is brought out of the coma; it's unclear when that will be.
Family members have set up an online fundraiser on CaringBridge.com called Blessings for Braxton to help the Bigelows with travel and medical expenses in the coming months. About $3,500 toward a goal of $10,000 has been raised so far.
Katy Moeller: 208-377-6413, @KatyMoeller
Tamarack Resort plans to have condos available before this year's ski season
Gorongosa girls club helps educate and empower future generations of Mozambique
Despite growth, Boise's crime rate drops to 25-year low. But drug crimes are on the rise.
By Katy Moeller
Boise is keeps growing, but the crime rate isn't. City officials touted an overall crime rate that has been falling for nearly 25 years in a press release Monday.
This women's clothing chain goes bankrupt, will close Treasure Valley store after sale
Business Columns & Blogs
Economists know the way to a better climate, and it's not the Green New Deal
Payette County wildfire contained after blaze burns across 2,000 acres, authorities say
28 Idaho legislators sign letter asking Boise State to stop supporting diversity programs
A people ravaged by civil war found new life in Idahoan's team's effort to plant coffee

redpajama_set_name: RedPajamaCommonCrawl


text:
From: Paris
Museums in Port Moresby
papuanewguinea.travel
National Museum and Art Gallery
Wander through the awe-inspiring collections at the National Museum & Art Gallery (NMAG). Here you will see a careful curation of objects from 19 provinces of Papua New Guinea. The earliest collection is by Sir William MacGregor and dates back to the 1800s during early administration of Papua. NMAG was built on Independence Hill in 1975 and opened to the public 1977. It is the national centre for anthropology, archaeology, natural history, contemporary arts, research and conservation.
Museums Museums in Port Moresby Art and Culture Art and Culture in Port Moresby Art Galleries Art Galleries in Port Moresby
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queensland.com
Museum Of Tropical Queensland
The award-winning Museum of Tropical Queensland provides a snapshot of this diverse region of North Queensland, from World Heritage listed rainforest and reefs to the story of the shipwrecked HMS Pandora, the ship sent to capture the Bounty mutineers. The Museum's family-friendly exhibitions and displays explore life in the tropics from prehistoric times to modern day. Located in the heart of Townsville, the Museum has temporary and permanent exhibitions to captivate visitors and their school holiday programs offer something for kids of all ages.
Museums Museums in Townsville
mackayregion.com
Artspace Mackay
Discover the arts and culture of the Mackay region when you visit Artspace Mackay, the regional art gallery and museum in the centre of the city. This architecturally award winning building provides visitors with an opportunity to learn about the people and history of Mackay through the Spirit and Place: Mementos of Mackay exhibition. Group tours can be booked with well trained volunteer guides who will provide visitors with insights into the exhibitions on display.
Museums Museums in Mackay Art and Culture Art and Culture in Mackay Learning Centers Learning Centers in Mackay Art Galleries Art Galleries in Mackay Activities for Kids Activities for Kids in Mackay
solomon.spla.pro
Solomon Islands National Museum
The Solomon Islands National Museum was first opened on June 1969. It is situated in Honiara, the capital, in Guadalcanal Province. It is the only national museum in the whole country at present. Solomon Islands National Museum is a museum, cultural centre, government department or ministry and also houses ww2 relics. The major part of the collection consists of cultural materials with some examples of natural history specimens, World War II relics and archaeological material. The Museum also houses a collection of Audio-Visual material and reference books. Number of items in the collection: over 2,000 items The collection contains items of the following types: Art, Heritage, Photography, Archaeological, Natural History (Animal Specimens, wet and dry), Contemporary, Scientific or Technology, Audio/visual material, Books or Journals, Posters, Pamphlets, Documents or Paper-based items, Outdoor sculpture, art or monuments.
Museums Museums in Honiara Art and Culture Art and Culture in Honiara
gladstoneregion.info
Gladstone Maritime Museum
Gladstone Maritime Museum focusses its collection on maritime history of the Gladstone Region. Visit to view history from prior to Captain Cook up to modern times. Special features include the shipwreck wall, Jenny Lind figurehead and a library. Naval and sailing ship models are of interest.
Museums Museums in Gladstone
Gladstone Regional Art Gallery & Museum
The Gladstone Regional Art Gallery and Museum was established in 1985 and is a community cultural initiative funded by the Gladstone Regional Council. It is dedicated to promoting art and heritage for and by the communities of the Gladstone Region and Central Queensland, acting as a focal point for the preservation and display of the region's history and cultural heritage.
Museums Museums in Gladstone Art and Culture Art and Culture in Gladstone
northernterritory.com
The Darwin Military Museum is located at East Point adjacent to the Defence of Darwin Experience. The site is within a heritage listed area that contains WWII fortifications. The No.2 gun turret is located within the Military Museum and, along with the No.1 gun turret, is one of the major fortifications at East Point. The Museum aims to collect artefacts from all conflicts where Australian troops were deployed
Museums Museums in Darwin
visitbrisbane.com.au
Queensland Gallery of Modern Art
Gallery of Modern Art (QAGOMA) is located across two riverside buildings in South Bank's Cultural Precinct. QAGOMA presents an evolving program of Australian and international exhibitions, with a focus on the contemporary art of Australia, Asia and the Pacific. Immerse the family in creativity at the Children's Art Centre and see the best in international film and video at the Australian Cinémathèque. QAGOMA offers cafes, modern dining and shopping to complete your visit.
Museums Museums in Brisbane Art and Culture Art and Culture in Brisbane Art Galleries Art Galleries in Brisbane Activities for Kids Activities for Kids in Brisbane
Queensland Museum & Sciencentre
Queensland Museum is the State's centre for natural history, cultural heritage, science and human achievement. Home to permanent and changing exhibitions and collections, the museum also provides innovative public programs, educational experiences plus holiday and early child hood activities. The Queensland Museum is also home to the Sciencentre, where visitors can take part in exciting - and educational - kinetic and interactive displays and experiments.
Museums Museums in Brisbane Art and Culture Art and Culture in Brisbane
discovercentralaustralia.com
The Alice Springs Telegraph Station is a historic museum precinct, presenting the story of the connection of Australia to the rest of the world through Telegraph Communication in 1871. Since being declared protected as a Historical Reserve in 1963, it has become the best preserved Station along the Overland Telegraph Line. Guided tours of the heritage precinct run from March to November at 9:30am, 11:30am, 1:30pm & 3:30pm. Visit the original "Alice Spring" and learn about the origins of the town. Enjoy NT roasted coffee and homemade snacks at the Trail Station Wi-Fi Cafe and browse through the gift and souvenir shop and send a postcard from the towns original red Postbox.
Museums Museums in Alice Springs Historical Sites Historical Sites in Alice Springs
The Museum of Central Australia
Central Australia's unique natural and geological history is explained at the Museum of Central Australia in Alice Springs. The evolution of the magnificent Central Australian landscape and the fascinating creatures that inhabited it, are described in the exhibitions. See meteorite fragments, fossils and interpretive displays as well as a replica of the ancient Alcoota waterhole. This internationally significant fossil site was home to giant freshwater crocodiles and the largest bird that ever lived. Contemporary Central Australia is explained in a display about the birds, mammals, reptiles and insects you will see as you tour the outback. The Museum of Central Australia also houses the Strehlow Research Centre, an important collection of film, sound, archival records and museum objects relating to Indigenous ceremonial life. The Strehlow Collection was accumulated by the Lutheran Pastor Carl Strehlow and his son Professor TGH Strehlow over two generations of anthropological research with the Aboriginal people of central Australia.
Museums Museums in Alice Springs Art and Culture Art and Culture in Alice Springs
surfersparadise.com
The Wax Museum
Prepare to be amazed as you experience the entertaining and unique wax museum right in the heart of Surfers Paradise. Meet people who made history, movie stars, scientists, explorers, villains, royalty and world leaders all presented with startling realism and dressed in authentically reproduced costumes.
Museums Museums in Gold Coast Family Fun Family Fun in Gold Coast
visitnsw.com
Lismore Historical Society & Museum
The award-winning Richmond River Historical Society Museum has one of the best historical collections in regional Australia. The museum is located in the heritage-listed former Lismore Municipal Building, with the main exhibition displayed within the old Council Chamber. Panels of local rainforest timbers line the walls of the museum.
Museums Museums in Gold Coast Art and Culture Art and Culture in Gold Coast
museenouvellecaledonie.nc
Musee de Nouvelle Caledonie
Built in the 1970s, the museum of New Caledonia asserts, in the late 1980s, its vocation as a museum of society and its objectives as to the enhancement of the Kanak civilization and the preservation of its heritage. Since then, this museum exhibits one of the most beautiful collections of Kanak art in the world, including monumental, and the most complete from the point of view of the themes it allows to address. In view of the Kanak collections of major European museums, the place occupied by the New Caledonia museum in the international museum world is particularly impressive.
Museums Museums in Noumea Art and Culture Art and Culture in Noumea Learning Centers Learning Centers in Noumea
newcaledonia.travel
Tjibaou Cultural Centre
Looking for the perfect introduction to local Kanak culture? Plan a visit to the Tjibaou Cultural Centre in Nouméa, where art, history, culture, knowledge and natural beauty combine to delight and educate adults and children alike. Housed within beautifully designed buildings and landscaped grounds, the centre sits just minutes from the city, offering a fantastic selection of permanent and temporary exhibitions to discover. A visit to the centre is a must for any Nouméa itinerary.
Museums Museums in Noumea Architecture Architecture in Noumea Art and Culture Art and Culture in Noumea Learning Centers Learning Centers in Noumea
Manning Entertainment Centre
Manning Entertainment Centre offer first class facilities for the performing arts, from school productions and amateur theatre to artists of national and international acclaim. The Manning Entertainment Centre is a 505 seat theatre in Taree, New South Wales and serves the residents of the Manning Valley, Great Lakes, Gloucester and Camden Haven regions.
Museums Museums in Port Macquarie Art and Culture Art and Culture in Port Macquarie
sydney.com
Australian Museum
Established in 1827, the Australian Museum (AM) is Australia's first museum and has been at the forefront of scientific research, collection and education for more than 185 years. Through exhibitions and other public programs the AM continues to inform and amaze generations of visitors about the unique flora, fauna and cultures of Australia and the Pacific.
Museums Museums in Sydney Art and Culture Art and Culture in Sydney
Smiths Expeditions
Bronte House
Bronte House's story begins in 1836 when William Mortimer Lewis, Colonial Architect, bought 42 acres of land at 'Nelson Bay' (the name given to the bay at Bronte Beach). He began building a house but when an econmic depression hit in 1843, was forced to sell the property before its completion.Robert Lowe, an English barrister and later NSW parliamentarian bought the property as a 'country residence' and finished the house in 1845. He and his wife Georgiana were some of the Bronte House's most charismatic inhabitants, despite only being in resident for four years.The house change hands quickly over the next couple of years, until the Ebsworth family bought the property in 1882. They were the longest private owners of Bronte House; the family occupying the property over three generations. In 1948 the Ebsworths sold the house and its ground to Waverley Council.
Museums Museums in Sydney Architecture Architecture in Sydney Art and Culture Art and Culture in Sydney
visitcanberra.com.au
Australia's stories come alive at the National Museum of Australia, on the shores of Canberra's Lake Burley Griffin. The Museum's stunning architecture provides an extraordinary place to explore the rich and diverse stories of Australia and its people.
Museums Museums in Canberra Architecture Architecture in Canberra
military-vehicle-museum.org.au
National Military Vehicle Museum
Welcome to the National Military Vehicle Museum, which is operated on a volunteer basis by members of the Military Vehicle Preservation Society of South Australia. The Society is one of many Military Vehicle Clubs across Australia and throughout the world and its members are dedicated to the preservation of vehicles that are of military origin. The museum was developed as a way of providing undercover storage for the vehicles and at the same time allowing the public to view them. The majority of the vehicles are privately owned by the Society members and therefore each member is responsible for their own vehicles. In addition the vehicles are not just restored to look at, they are rebuilt to authentic running condition and most are road registered and driven regularly. You will see military vehicles and items ranging from the First World War to current times.
Museums Museums in Adelaide
southaustralia.com
Port Adelaide Historical Society Museum
The Port's local community museum is located a few minutes from the centre of Port Adelaide and contains historic collections of Port Adelaide, Semaphore, Lefevre Peninsula and environs, including the maritime collection of Keith Leleu, which was commenced in 1962 to address the loss of all things maritime. It houses a wide collection of artefacts, published materials, books, photographs and models; arranged to interest all who love ships and the sea.
Migration Museum
Discover the history of South Australia through the stories of people and communities. The Migration Museum works towards the preservation, understanding and enjoyment of South Australia's diverse cultures. It is a place to discover the many identities of the people of South Australia through the stories of individuals and communities.
Museums Museums in Adelaide Art and Culture Art and Culture in Adelaide
communitywebs.org
The summer residence of South Australia's early governors, Old Government House, sits in the heart of Belair National Park. The house is an excellent example of Victorian architecture and is set amongst magnificent gardens. It was built on the Government farm as the former summer residence of several early Governors of South Australia. The cottage is an excellent example of Victorian-style architecture. It is looked after by a dedicated group of volunteers and the Friends of Old Government House in conjunction with DEWNR.
Museums Museums in Adelaide Architecture Architecture in Adelaide Historical Sites Historical Sites in Adelaide
poklat.com
Davao Museum and Ethnography
The evolution and richness of Philippine history ensembles a simplest framework of how social change significantly takes place to effectively initiate its highest form of realization.-Socio-cultural development.
Museums Museums in Davao
indonesia.travel
Balla Lompoa Museum
An impressive "Changing of the Guard" ceremony takes place at the Ballalompoa Museum, the former Palace of the Sultan of Gowa. Wearing resplendent bright red costumes, "soldiers" involved in the "Changing of the Guard" at the former Gowa Palace will perform their drill for the public to watch every morning at 09.0 am, promised the Bupati (Regent) of the district of Gowa. Built in 1936, the Palace of the Sultan of Gowa, formerly also known as the Sungguminasa Palace, is a large wooden structure built on stilts and has an imposing roofed staircase that leads to the reception room. It has five wooden panels in the gable of the roof, which denotes royalty and may be used only by the Sultan. The Ballalompoa Museum today houses the resplendent treasure and regalia of the former Sultan of Gowa, among which the Gowa gold crown, bracelets with precious stones, royal swords, keris and gold jewellery. Every year a special ceremony called "Acera Kalomponag" is held to cleanse and weigh the gold crown, which is believed to have been inherited from the first female ruler of Gowa, called I Tanisamanga. According to belief, the crown never weighs the same each year it is weighed. When its weight increases it is a good omen that all will be well in the region.
Museums Museums in Makassar Architecture Architecture in Makassar Art and Culture Art and Culture in Makassar
Fort Rotterdam
Standing majestically at the western coast of Makassar, South Sulawesi. Fort Rotterdam is recognized as the city's most iconic landmark. With historical traces dating back to the Kingdom of Gowa from the 16th century to colonization by the Dutch, this Fort has silently witnessed many episodes in Makassar's history, playing a most essential role in its development. Its magnificence and authenticity has always captivated those who set eyes on it. Originally called Benteng or Fort Jumpandang or Ujung Pandang, the huge complex was first built in 1545 in the era of Imanrigau Daeng Bonto Karaeng Lakiung or Karaeng Tunipalangga Ulaweng, the tenth King of Gowa. Initially, the fort was made from a mixture of Stone and burnt clay, and took the shape of a typical square Portuguese architectural style. The fort was also expanded and took on a new shape resembling a sea turtle, thus the fort gained a new name, Benteng Pannyua (Penyu) or Fort Sea turtle. The shape is not only unique, but also contains deep meaning. For just as a sea turtle lives both on land and at sea, the glory of the Gowa Kingdom also stretched on land as well as over the seas.
Museums Museums in Makassar Architecture Architecture in Makassar Historical Sites Historical Sites in Makassar
visitballarat.com.au
Art Gallery of Ballarat
The Art Gallery of Ballarat, Australia's oldest regional gallery, located in the heart of Ballarat's central heritage precinct, boasts a stunning collection of Australian art. Its busy exhibition schedule makes the Gallery an exciting and dynamic place presenting a range of exhibitions from cutting edge contemporary art to themed shows drawn from the Gallery's extensive collection, and special interest ticketed shows.
Museums Museums in Melbourne Art and Culture Art and Culture in Melbourne Art Galleries Art Galleries in Melbourne
melbourne.com.au
Chinese Museum Melbourne
A living part of Melbourne's modern Chinatown, the Chinese Museum is a national museum, brought into being in 1985 to document, preserve and display the history of Australians of Chinese descent who have helped shape and develop what it means to be Australian today. An exciting range of changing exhibitions, heritage tours and public seminars are some of the ways by which the Chinese Museum shares the past, culture, and values of Australia's Chinese community with the wider public. Educating existing and future generations of Australians it places importance on our identity and what is means to be Australian. A goldfield's mine re-creation, displays about early Chinese industry such as furniture making and the banana industry, imperial garments and images of the Young Chinese League's debutantes of the 1930s onward, evoke some of the richness of the Chinese contribution to Australia's history. The Museum is also home to Dai Loong and the Millennium Dragon, the largest dragon in the world, which weaves through the streets of Melbourne following a 100 year old Melbourne tradition. The Museum is a popular and important educative resource for teachers and educators alike visited by a huge number of school children 25,000 per annum to learn about multiculturalism, local and Australian history, Chinese arts and culture.
Museums Museums in Melbourne Art and Culture Art and Culture in Melbourne Learning Centers Learning Centers in Melbourne Activities for Kids Activities for Kids in Melbourne
ngv.vic.gov.au
The National Gallery of Victoria (NGV) is the oldest and most visited gallery in Australia. Situated over two magnificent buildings – NGV International and NGV Australia – the Gallery hosts a wide range of international and local artists, exhibitions, programs and events; from contemporary art to major international historic exhibitions, fashion and design, architecture, sound and dance.
Museums Museums in Melbourne Art and Culture Art and Culture in Melbourne Learning Centers Learning Centers in Melbourne Art Galleries Art Galleries in Melbourne Activities for Kids Activities for Kids in Melbourne
visitvictoria.com
Portarlington Mill
Located just 300 metres from the beach, the Portarlington Mill was constructed in 1856 with rough local sandstone and Baltic Pine floors. The interior spaces are quintessential mid 19th century industrial style: large open spaces with heavy timber posts and beams enclosed by natural stone. The Mill is one of the few remaining and preserved Victorian flour mills.
Museums Museums in Melbourne Architecture Architecture in Melbourne Historical Sites Historical Sites in Melbourne
National Wool Museum
The National Wool Museum is housed in a beautifully restored 1872 bluestone wool store close to the vibrant new Waterfront Geelong. The Museum is Australia's largest comprehensive museum of wool, showcasing the region's wool story from the sheep's back to the clothes rack and from the birth of the industry in the 1840's to its place in the world today.
Museums Museums in Melbourne
Old Geelong Gaol
The Old Geelong Gaol and Museum has a great variety of displays that tell the history of the gaol and visitors get hands on experience of the harsh conditions prisoners and staff endured at the facility. Visitors can experience solitary confinement cells, get a taste of prison life, meet some infamous figures who have spent time at the gaol and see the original hallows.
Queenscliffe Maritime Museum
Sea, sand and sail are the key to Queenscliff's history. Visit Queenscliffe Maritime Museum to discover the treasures of its rich maritime heritage, shaped by its proximity to the entrance of Port Phillip and its notorious 'rip'. From sea pilots to sailors, fishermen to boat builders, light keepers or ferry captains, the sea has created a world full of stories of boats and maritime industry.
Queenscliffe Historical Museum
The Queenscliffe Historical Museum is situated in Hesse Street between the post office and library, in the township of Queenscliff. Opened in 1974, specifically to house socially historical materials peculiar to the Borough of Queenscliffe (which includes Queenscliff, Point Lonsdale and Swan Island) the museum is home to many thousands of items including photographs, documents, paintings, newspapers and toys, just to mention a few.


text:
Editor Pick Features, Features, Side Feature
A Conversation with Colby Republicans
October 30, 2019 10:15 pm	 · 0 commentsViews: 244
By: Claire Borecki on October 30, 2019.
Two members of the Colby Republicans sat down with the Echo to discuss their perspectives, hopes, thoughts, and anxieties about former governor Paul LePage's visit to campus, as well as the club's plans for the future.
Meredith Allen `21 of Dover, Massachusetts, is studying economics and data science. On campus she is the co-president of Colby Republicans, an officer and choreographer for Colby Dancers, on the leadership board of the Goldfarb student engagement committee, and a CCAK mentor. When asked what she wants to do after Colby, she shakes her head sheepishly. "Something…in business," she said. "Don't really know."
Her fellow Colby Republican and close friend, Ellie Harlan `21 is from Westwood Massachusetts. On campus she is the secretary of the Colby Republicans, is also on the Goldfarb student engagement committee, in the intervarsity Christian fellowship, and in Colby Dancers. She's a government major, and she wants to go into the state department in some capacity, "preferably in the Middle East," she said. She's taking Arabic and studying Middle Eastern politics.
The members of the Colby Republicans go to a meeting in Augusta every month with the Maine State Republicans, a meeting LePage attends frequently. It was here that LePage offered to visit college campuses and share his experiences; he personally approached the Colby Republicans faculty advisor, Professor Dave Findlay, and offered to come speak at Colby—at no cost to the College.
"When Findlay told us about their [Findlay and LePage's] conversation we then extended an invitation to him, given that he was willing to come for free and share his experiences with us," Allen said.
Allen and Harlan both agreed that this was slightly unusual, but free is free, and it was an opportunity. "We don't often get a chance to connect with a lot of higher profile figures in the state. We're a small club," Harlan said.
During the talk, LePage introduced himself and told his story. "A lot of people aren't aware [of his story]," Allen noted. LePage discussed several issues. He only spoke for about twenty minutes, and then opened the floor for questions from the audience. Apparently, this was somewhat by design.
"He really wanted it to be mostly question and answer," Allen relayed. "Which we thought was a really good idea," Allen smiled, "because we weren't giving him a place to just[…]spew[…]anything," she says. "We really wanted it to be targeted at what students wanted to hear from him and give them the opportunity to challenge him on things that he's done and said."
"I think it went well," Harlan explained. Many students and faculty asked questions, and a number of local Mainers who attended the event were able to voice their questions as well. "As a club we wanted to have a dialogue. We wanted to create a space where we could invite a conservative speaker and allow[…]that back and forth I don't think we see a lot on campus," Harlan stated. "And in that way I think we got what we wanted out of it." LePage also stayed after his speech to talk to students individually, remaining much longer than his allotted speaking time, seemingly enthused about talking to students.
Allen and Harlan had to learn a lot along the way, particularly about the process of bringing in a very high-profile speaker. "We learned a lot[…]about allowing both sides to express their feelings about it," Allen said. The Colby administration was supportive of this, and worked with the republicans to create standards for the event, and guidelines for what to do "if anything happened," Allen said. "I think it's progress such that we could bring more speakers in the future, whether they're controversial in the same way…or…not," Allen laughed, "which might be preferable." There were laughs all around at that. "But I do think most Republicans would face some controversy on campus."
When asked, with an amount of preemptive disbelief, if they agreed with all of LePage's views, Harlan responded quickly: "absolutely not."
They did provide some defense for his invitation. "We shouldn't have to justify bringing him more than him being governor for eight years," Allen opined. "His comments don't delegitimize what he did. He should still have a platform to share what he did, and we weren't giving him a platform to share what we don't endorse."
"I think he did a lot of good for the state, and what he's said should not preclude him from being able to speak to his experience in politics and his ideas for how Maine can be better," Harlan added. "But I think because of what he has said that was why we wanted to have the opportunity for students to challenge him. I think they should, he's said some really…" Harlan trailed off, but the point was made.
When asked if LePage is racist, Harlan said "I don't know his heart."
When asked about how LePage's comments, particularly surrounding race, could threaten the identities of students on campus, and what they thought of the idea that LePage shouldn't be given any platform to speak at all, Harlan said, "We didn't want to give him an opportunity to say anything [that would threaten anyone]."
"We were really concerned about that," Allen shared. When asked about giving LePage a platform for hate speech, both reply "absolutely not."
"We told him, 'we would like you to talk about the economy, your background, your experience in politics,' and we were very clear with him in the many correspondences that we had what the parameters were. We didn't want students in the audience to feel…to hear him say something that would be hurtful."
In addition to concerns about hurtful comments, the two republicans had some thoughts about the wider political world.
"You should be prepared to interact with people of all different kinds," Allen asserted. "The world is not a safe space."
Author: Claire Borecki