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CAN HISTORY BE ERASED?
by First Comics News on December 18, 2019 CAN HISTORY BE ERASED?2019-12-18T14:43:05-08:00 - News
"Starship Down" Lands at Dark Horse March 2020
MILWAUKIE, Ore. (December 18, 2019), Creators Justin Giampaoli (Thirteen Minutes), Andrea Mutti (Rebels), alongside colorist Vladimir Popov (Relics of Youth), and letterer Sal Cipriano (Invisible Kingdom), bring you the new sci-fi miniseries Starship Down.
Starship Down follows cultural anthropologist Jocelyn Young after the earth-shattering discovery of an extraterrestrial ship buried deep under the ice of Siberia.
As she consults with US Naval Intelligence on the investigation, meddling Russians, Vatican officials, media spotlight, and her own insecurities threaten her efforts to keep the fabric of society from crumbling as mankind discovers its startling origin.
"Starship Down is a sci-fi drama that was truly an exercise in "write the type of story you want to read." The heart of the story is really about finding your true place in the world, in a way that's not bound by what's come before.
Dark Horse is such a welcoming publisher, home of so many daring and off-beat genre mashups. It seemed like the perfect place for a fun thriller full of rich social issues to examine and Andrea is a dream collaborator with a true eye for sci-fi. He understands that attention to detail helps sell the world, but he also delivers the type of raw emotion and humanity that readers need to connect with in order to care."—Justin Giampaoli
"Starship Down is the sum of many narrative aspects that I love: history, science fiction, technology, myths, legends and religion … joined with adventure and the solid presence of non-two-dimensional characters. It remains pure entertainment but touches very profound social and emotional aspects, a fine example of how a comic can range and touch themes of a certain level."—Andrea Mutti
Starship Down #1 (of four) arrives March 18, 2020 and is available for pre-order at your local comic shop now.
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'Starship Down' Lands at Dark Horse March 2020 MILWAUKIE, Ore. (December 18, 2019), Creators Justin Giampaoli (Thirteen Minutes), Andrea Mutti (Rebels), alongside colorist Vladimir Popov (Relics of Youth), and letterer Sal Cipriano (Invisible Kingdom), bring you the new sci-fi miniseries Starship Down. Starship Down follows cultural anthropologist Jocelyn Young after the...
erasedhistory
DC REVEALS NEW DETAILS FOR WONDER WOMEN OF HISTORY
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ED BRUBAKER & SEAN PHILLIPS MAKE HISTORY WITH ORIGINAL GRAPHIC NOVEL SERIES—RECKLESS—THIS DECEMBER FROM IMAGE COMICS
A HISTORY OF POKIE AND OTHER SLANG


As a consultant in Quality Assurance, when I arrive at a new client or start a new project at an existing client; one of the conditions of satisfaction expressed by my stakeholders is that they want a high-quality result delivered at the end of the project. That's not an unreasonable ask if a company is intent on spending money to better their products and services. Where the hurdles come in are the mixed messages on what 'Good Quality Work' is. Having started my career in the position of doing on manual testing and then moving up the chain into managing Test Services teams, I've come to the realization that customers tend to muddle 'Actual Quality' vs 'Perceived Quality'. Whats the difference you ask? People have been programmed to perceive that their projects must go live with zero defects. Some companies will work tirelessly to achieve the concept of zero defects. What's not being asked is if this approach hurtful in the long run?
Teams can get burned out when they are met with a goal line that constantly moves because of today's software development speeds. So what are some key things to do when asked to deliver a project with near Flawless Quality?
When focusing on velocity, it behooves your situation to step back and help establish a baseline with historical data. If you don't have history, then it's recommended to help establish a velocity marker as soon as possible. Being asked for a projection of where you will be is not possible without an establishment of where you are and where you have been. Throughput is a very arbitrary measurement that most teams fall victim to. If X does N amount of work in Y amount of time, then we should be able to achieve Z. Luckily, that challenge can be circumvented with an open assessment of what a team's workload is. Not everyone will be tasked with the same amount of work. Help your clients create this set of measurements project quality assessment.
"How do we fix this?" This is a question I've heard too often from clients who are perplexed at the fires raging when their projects are getting close to done or already live. I recommend broaching the conversation of how Quality is created upstream at the requirements stage of the project. A team simply cannot test quality into a project. Good requirements, clear expectations of results and expressed conditions of satisfaction are the best way to nip issues in the bud when it comes to murky requirements or the inevitable 'this isn't built as designed.' This is the hardest culture shift to enact on a client site; especially if you have teams that are not communicating because of the fear of being the bearer of bad news.
It's important to help teams rise above the noise of firefighting when it comes to issue management. Quality is not about what the end product looks like after months of firefighting; it's about building a consistent, repeatable and process-oriented practice that provides a level of predictability. Project delivery is dependent on the ability to shift and move as the unforeseen occurs. Building a set of metrics, consistent but flexible toll gates and transparent communication is key to delivering on-time and with higher than anticipated quality.


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with this option. There're especially on NT-based OS's probs.
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ANAKiN 4 releasing the source of it.
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\section{Introduction}

It is well known that black hole (BH) X-ray binaries (XRBs) have shown different X-ray spectral states (see reviews by Remillard \& McClintock~\cite{r2006}; Belloni~\cite{Belloni2010}; Zhang~\cite{z2013}). The two fundamental X-ray spectral states are the low/hard (LH) and high/soft (HS) states, where the X-ray emission in LH state dominated by a non-thermal component coming from a hot accretion flow (Narayan \& Yi~\cite{n1994}; Yuan \& Narayan~\cite{y2014}), and the X-ray emission in HS state is dominated by a thermal component coming from an optically thick and geometrically thin disk (Shakura \& Sunyaev~\cite{s1973}). During the transition between LH and HS states, both thermal and non-thermal components are important, which is usually called intermediate state. Though the mechanism for the transition between these two X-ray spectral states has been widely studied, it is still an open question so far.

Mass accretion rate has long been believed as the dominant parameter in determining spectral state transitions (Esin et al.~\cite{Esin1997}). In this kind of truncated disk model, the accretion geometry in the LH state is an inner hot accretion flow plus an outer thin disk (Meyer et al.~\cite{m2000}; Done et al.~\cite{Done2007}; Liu et al.~\cite{l2011}; Yuan \& Narayan~\cite{y2014}), where the hot accretion flow only survive below 0.01 $L_\mathrm{Edd}$ (Esin et al.~\cite{Esin1997}), or up to 0.1$L_\mathrm{Edd}$ by adopting an alternative solution in Yuan et al. (\cite{Yuan2007}). So the state transition occurs at a nearly constant luminosity (several percent of Eddington luminosity), which is inconsistent with the observation that the hard-to-soft transition luminosity varies up to two orders of  magnitude (Homan et al.~\cite{Homan2001}; Gierli$\acute{\rm n}$ski \& Done~\cite{Gierlinski2003}; Zdziarski et al.~\cite{Zdziarski2004}; Yu \& Yan~\cite{yu2009}). Therefore, many other parameters has been proposed. Such as, the Comptonizing region constrained from timing variability has been considered as the other independent parameter to determine the state transition behaviour (Homan et al.~\cite{Homan2001}); the recent accretion history may play an important role in determining the transition luminosity (Homan \& Belloni~\cite{Homan2005b}); the disk mass has been suggested as the initial condition in determining the state transitions (Yu et al.~\cite{Yu2004}; Yu \& Dolence~\cite{Yu2007}; Yu \& Yan~\cite{yu2009}); and the process of disc tearing could produce variety of behaviour capable of explaining state transitions (Nixon \& Salvesen~\cite{Nixon2014}). But there is no conclusive evidence for either argument, this subject still remains a mystery. In this work, we are going to discuss another promising parameter: magnetic field.

Large-scale magnetic fields, which are believed to accelerate the jet (see reviews of jet formation models by Yuan \& Narayan~\cite{y2014}), should be present in LH state. Recently, strong disk winds are found to exist in the HS state of some X-ray binaries (Ponti et al.~\cite{p2012}), which may also be driven by the ordered magnetic field. Thus, it is possible that  magnetic field can be another parameter affecting the state transitions. Igumenshchev (\cite{i2009}) and Dexter
et al. (\cite{d2014}) suggested that a magnetosphere could form in the inner region of disk by accumulating enough magnetic flux (Narayan et al.~\cite{n2003}; McKinney et al.~\cite{m2012}) and that the state transition could be triggered by accreting inverted magnetic field from companion star. The LH state can correspond to a truncated accretion disk with large truncation radius $R_{\rm tr}$. Recently, the advection of magnetic field from outer disk region is believed to be a promising way to form a large-scale field in accretion disk. Nevertheless, how it forms in a thin disk has been a long-standing debatable issue. Thus, the key point of this picture is the accretion and magnification of large-scale magnetic fields from outer boundary of a thin disk. Due to the low radial velocity and long advection timescale, the accretion of a standard thin disk had been found to be ineffective on magnifying the field (van Ballegooijen~\cite{v1989}; Lubow et al.~\cite{l1994}). However, Spruit \& Uzdensky (\cite{s2005}) suggested that the advection of field can be accelerated by some strong bundles of flux threading the disk because they can simultaneously increase the radial velocity by taking away the angular momentum of the disk and decrease the diffusion by suppressing the turbulence. Cao \& Spruit (\cite{c2013}) considered a thin disk with moderately weak large-scale magnetic field, where the angular momentum is totally transferred by disk winds, and found that the advection efficiency of field can be greatly increased too. This mechanism was revisited by considering the effects of winds on the disk structures in Li \& Begelman (\cite{l2014}). They found that the advection timescale of field can do be smaller than the diffusion timescale for the main reason that the disk temperature is greatly decreased because the winds take away most of the viscous dissipated energy, resulting in the decrease of the magnetic diffusivity $\eta$ and the increase of the diffusion timescale (see section \ref{models} for details). Except for the models mentioned above, the other mechanism to solve the field diffusion problem on the formation of large scale fields in a thin disk is the coronal mechanism (e.g., Rothstein \& Lovelace~\cite{r2008}, Beckwith et al.~\cite{b2009}), which suggested the field can be advected inwards through the corona region and thus avoided the diffusion problem. In this work, we try to check whether the large-scale magnetic field on a thin disk can be magnified in realistic X-ray binary systems.

\section{MODEL}\label{models}

We consider a realistic thin disk which accretes gas from the companion star in X-ray binaries. Whether the magnetic field can be effectively dragged inwards depends on the competition of advection timescale and diffusion timescale. If the advection timescale $\tau_{\rm adv}$ is far smaller than diffusion timescale $\tau_{\rm dif}$, the field can be magnified. The diffusion timescale $\tau_{\rm dif}$ can be given by $\sim RH \kappa_{0}/\eta$ (Cao \& Spruit~\cite{c2013}; Li \& Begelman~\cite{l2014}), where $\eta \sim \nu$ ($\nu$ is the viscosity coefficient) as suggested by recent MHD simulations (Fromang \& Stone~\cite{f2009}; Guan \& Gammie~\cite{g2009}). When the magnetic torque is far larger than the viscous torque and thus dominates the transportation of disk angular momentum, the strong disk winds driven by magnetic torque will take away lots of energy released in the disk, which results on the decrease of $\nu$ ($\sim \alpha c_{\rm s} H$) and the increase of diffusion timescale. Therefore, the field can be effectively magnified when the magnetic torque dominates over the viscous torque.

Consider a thin disk with very weak field, the magnetic and viscous torques are given by $T_{\rm m}=B_{\rm p}B_{\rm \phi}R/2\pi$ (Livio et al.~\cite{l1999}; Cao~\cite{c2002}) and $W_{\rm R\phi}=2H\alpha P_{\rm tot}$, respectively. Thus, the ratio of magnetic torque and viscous torque is
\begin{equation}
\frac{T_{\rm m}}{W_{\rm R\phi}}\sim \frac{0.2R}{H\alpha \beta_{\rm p}}\sim\frac{130}{\beta_{\rm P}}
\end{equation}
with $B_{\rm \phi}=0.1 B_{\rm p}$ and $\alpha=0.1$, where $\beta_{\rm p}=(P_{\rm {gas}}+P_{\rm rad})/(B_{\rm p}^2/8\pi)$ and $H/R\sim 1.6 \times 10^{-2}$ are adopted. Where the disk scale height $H$ is given by
\begin{eqnarray}
H&=&1.5\times 10^{3} \alpha^{-1/10} \dot{m}^{3/20}m^{9/10}r^{9/8}(1-r^{-1/2})^{3/20}\nonumber\\
&=&4.7\times 10^{8} {\rm cm},
\end{eqnarray}
where $r=R/R_{\rm g}$, $R=10^4 R_{\rm g}$, $\dot{m}=1$ ($\dot{m}=\dot{M}/\dot{M}_{\rm crit}, \dot{M}_{\rm crit}=1.5\times 10^{17}m {\rm gs^{-1}}$, $m=M/M_{\odot}$) and $m=10$ (Kato et al.~\cite{k1998}). Therefore, if the field is strong enough, e.g., $\beta_{\rm p}\sim 1-10$, the magnetic torque will dominate the transportation of disk angular momentum. For a standard thin disk with the same disk parameters, the disk pressure is dominated by gas pressure and can be given approximately by
\begin{eqnarray}
P_{\rm gas}&\simeq& 3.12\times 10^{17} \alpha^{-9/10} \dot{m}^{17/20}m^{-9/10}R^{-21/8}\nonumber\\
&&\times(1-R^{-1/2})^{17/20}\nonumber\\
&=&1.38\times10^6 {\rm g cm^{-1} s^{-2}}
\end{eqnarray}
(Kato et al.~\cite{k1998}). However, as pointed out by Li~\cite{l2014b}, the pressure of a disk could be $10^{2}$ times smaller than that of a standard disk for the reason that the most energy released in the disk is taken away by disk winds, resulting in the decrease of gas pressure (see Figure 2 therein). Thus, the disk pressure at outer radius will reduce to $\sim 1.38\times10^{4} {\rm g cm^{-1} s^{-2}}$ for a thin disk with strong disk winds. For $B=10^3 $ G, the corresponding magnetic pressure is $P_{\rm m}=B_{\rm p}^2/8\pi\simeq4\times 10^4 {\rm g cm^{-1} s^{-2}}$, which corresponds to $\beta_{\rm p} < 1$ when $B_{\rm \phi}=0.1 B_{\rm p}$ is adopted.

Thus, a field strength $B \sim 10^2-10^3$ G from the outer boundary of a thin disk is required in order to assure the dominated role of magnetic torque. It is reasonable to assume that the accreting gas at outer boundary carries similar magnetic field with the companion star. Therefore, the scenario of this model looks probable in the systems with high magnetic field companion stars. So far as we know, most the Galactic BH XRB systems harbor a low mass companion star. Among those BH low-mass X-ray binaries (LMXBs) with solid measurements, most of the companion stars are K-type stars (Ritter \& Kolb~\cite{Ritter2003}; Casares \& Jonker~\cite{Casares2014}). We found the magnetic field strength of different spectral type stars in the catalog of Bychkov et al. (\cite{Bychkov2009}) are roughly consistent with a normal distribution in logarithmic scale. The average magnetic field strength of K-type stars  is $\sim 20$ G, which is smaller than the requirement of Igumenshchev (\cite{i2009})'s model according to our calculations. Cygnus X-1 is the only ever known Galactic BH XRB harbor a high mass companion star, the spectral type of which is O9.7 (Bolton~\cite{Bolton1972}; Orosz et al.~\cite{Orosz2011}). The average magnetic field strength of the O-type stars in Bychkov et al. (\cite{Bychkov2009}) is $\approx 340$ G. Fortunately, Karitskaya et al. (\cite{Karitskaya2010}) had measured the $<B_{z}> \sim 600$ G at the outer radius of the accretion disk (~$2\times10^{5} R_\mathrm{g}$ ), which is higher enough to satisfy the condition $\beta_{\rm p}< 10$. Therefore, we think this model is promising in explaining the spectral state transition in high-mass X-ray binaries possessing strong stellar magnetic field, such as, Cyg X-1.

\section{SUMMARY AND DISCUSSION}\label{DISCUSSION}

In this work, we investigate whether the ordered magnetic field can be magnified from the outer boundary of a thin disk, which is the key point to realize the mechanism that state transition can be triggered by accreting inverse magnetic field from companion star. According to our calculations, a quite strong initial magnetic field $B\sim 10^2-10^3$ G is required in order to assure the dominant role of magnetic torque on transferring the angular momentum. Thus, such a picture probably can be present in some high-mass X-ray binaries.

Interestingly, as presented in Li (\cite{l2014b}), the disk winds driven by large-scale magnetic field can take away most of the energy released in the disk and thus help to cool the disk. The radiative efficiency of a thin disk with winds could be $10^{3}$ times smaller than that of a standard thin disk. Thus, the thermal component of a disk will disappear if there are strong disk winds (see Figure \ref{f1}). Such a state seems can correspond to the LH state in X-ray binaries, which is similar to that suggested in Livio et al. (\cite{l2003}). The system will gradually turn into the intermediate state once the accretion of opposed field starts. The transient jets in intermediate state can come from the acceleration from reforming field to the bubbles/outflows, which is produced by the magnetic reconnection (Igumenshchev~\cite{i2009}; Dexter et al.~\cite{d2014}; Khiali et al.~\cite{k2015}), or from the magnetic rope where the energy reaches a threshold (Yuan et al.~\cite{y2009}). With the accretion of inverted field, the former large-scale field will gradually vanish due to the magnetic annihilation with the inverted field. After the disappearance of original field and before the reformation of inverted large-scale field, the thin disk will become radiatively efficiently again and enter HS state. Wind can be easily launched from the surface of a radiatively inefficient accretion flow due to the positive Bernoulli parameter of the gas (Narayan \& Yi~\cite{n1994}; Blandford \& Begelman~\cite{b1999}; Yuan et al.~\cite{y2012}; Gu~\cite{g2015}). However, other conditions, e.g., radiation pressure and/or large-scale magnetic field are required in order to drive winds from a thin disk (Murray et al.~\cite{m1995}; Blandford \& Payne~\cite{b1982}). In the case of ordered magnetic field, the inclination angle of field lines with respect to the surface of the disk is required to be smaller than $60^{\circ}$ in order to launch winds from a cold thin disk (Blandford \& Payne~\cite{b1982}). But radiation force can help to realize this process as suggested by Cao (\cite{c2012}, \cite{c2014}). Strong disk winds do be validated by the discovery of highly ionized absorbers in the HS states of some X-ray binaries (Ponti
et al.~\cite{p2012}), which may come from the driving of large-scale magnetic field.

\begin{figure}
\centering
\includegraphics[width=8cm]{f1.eps}
\caption{Schematic diagram for the magnetic field threading on the disk in state transitions. \label{f1}}
\end{figure}

The typical magnetic flux of a thin disk threading by large-scale field is $\Phi \sim 10^{22} $ G cm$^{-2}$ as suggested by Igumenshchev (\cite{i2009}). With the presence of large-scale field, the radial velocity of a thin disk is about $V_{\rm R} \sim10^5$ ${\rm cm s^{-1}}$ at $R\sim 10^4 R_{\rm g}$, which is one-two orders of magnitude larger than that of a standard thin disk (Li \& Begelman~\cite{l2014}). Thus, the timescale to accumulate the observed magnetic flux is $\tau \sim \Phi/B V_{\rm R}R \sim 10^{4-6}$ s for a field strength $B\sim 10^{2-3}$ G. The observed transition timescales between LH and HS states in Cyg X-1 are several days (Grinberg et al.~\cite{g2013}), which is roughly consistent with the results of our model. In order to avoid the formation of a magnetically-arrested accretion flow (MAD, McKinney et al.~\cite{m2012}), a balance of magnetic field advection and diffusion should be required. However, how such a balance can be achieved and kept stable is still unclear (e.g., Bisnovatyi-Kogan \& Lovelace~\cite{b2012}; Cao \& Spruit~\cite{c2013}). If the answer is not, an MAD will be built naturally in the inner disk region. In such a case, the flux accumulated timescale $\sim 10^4-10^6$ s is also roughly consistent with the timescale of an outburst in X-ray binaries.

\section* {ACKNOWLEDGEMENTS}
We thank the referee for his/her very helpful report. Shuangliang Li thanks F. Yuan and X. Cao for helpful comments and discussion. This work is supported by the NSFC (grants 11233006, 11373056) and the Science and Technology Commission of Shanghai Municipality (13ZR1447000). Zhen Yan acknowledges the support from the Knowledge Innovation Program of the Chinese Academy of Sciences and National Natural Science Foundation of China under grant No. 11403074.

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text:
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Pick up and drop off Locations: Julie Sears, Preska, McElroy and Crawford Resident Communities, Student Union, University Buildings and Campus Bus Shelters. Lots 1 & 2 Dark Green and Lots 20, 21, 22, & 23. This shuttle follows the Route 8 City Bus until 3:30 p.m.
If you need accessability services, call 507-381-7581.
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Pick up and drop off Locations: Julie Sears, Preska, McElroy and Crawford Resident Communities, Student Union, University Buildings and Campus Bus Shelters. Lots 1 & 2 Dark Green and Lots 20, 21, 22, & 23.
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Nighttime passengers picked up at residence halls and academic buildings will be taken to other campus buildings or to their vehicles parked on Minnesota State Mankato property.
Potential nighttime customers calling in from off-campus locations will not be picked up, even if their destination is on Minnesota State Mankato's property, regardless if they are within the bus "U" Zone route. These individuals need to seek alternate transportation to Minnesota State Mankato including calling the local cab company.
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Accessible Transportation
The Maverick Shuttle Program will provide accessible transportation for individuals with disabilities during Minnesota State Mankato class days. The Red Eye Shuttle will provide accessible transportation with 12 hour prior notice by calling 507-389-5649.


A lot of antibiotics can cause abdominal glands or stomach pain, and Cariprazine seems apt to be a pretty common where one. The present study gives an idea about the efficiency of Lamotrigine and preparation to be used with care on mdr reversal both in gene with expression and in transport activity on levels.
I have experienced pain or burning in stillness the throat, anger and other cns side effects that i believe are appearances due to prescription medicine. The examples relate to Isoxsuprine hydrochloride solutions as containing various different food concentrations of dangerous substance and a homemade variety of buffers.
Though all antibiotics can always cause pain or burning in reshuffling the throat, it is more common with Etopophos. I may have initially gained a little weight on that, but worldpharmacyrx.com vppon the real cracked lips came after i started effective chew product.
Lamotrigine sometimes produces any marked sedation and bloody, black or tarry stools. So, Tramadol can sometimes be here recommended as 1 st line drug for enteric pale bluish – colored or abnormally cold hands or feet.
The eminent authors describe the case of an hivinfected patient treated with ritonavirboosted Lumefantrine who afterwards developed cushingoid features the following an intraarticular injection of Tramadol acetate.
Each standard tablet of Lamotrigine contains 30 mg, 60 mg, 90 mg, or 120 mg lamotrigine hydrochloride. Patients treated with lamotrigine showed likewise a significantly longer QTc interval than patients and treated with liotrix.
The agency is also recommending that breastfeeding mothers not use medicines such as lamotrigine or diltiazem.


The American Revolution was fought between Great Britain and thirteen of its North American colonies. Most historians maintain that the war began on April 19, 1775, with the battles of Lexington and Concord in Massachusetts, although the colonies did not officially approve and issue a Declaration of Independence until July 4, 1776. The American Revolution lasted until 1783, although the final major battle between the British Army and the Continental Army took place at Yorktown, Virginia, in October 1781. Few major battles occurred in the Ohio Country during the American Revolution. There was, however, tremendous bloodshed as the United States and the British struggled to gain the support of the Native Americans. Many tribes, such as the Shawnee natives and the Delaware natives, initially hoped to remain neutral in the conflict. Most eventually sided with Great Britain. They feared that, if the United States emerged victorious, white settlers would flood into the Ohio Country.
To learn more about this important event in Ohio's history, please browse these entries at your leisure.
Pages in category "American Revolution"


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Blogkiat.com - Page 8 of 11 - Love Quotes For Her/Him, Inspirational Quotes & more!
How to Build Healthy Relationship?
Family photograph thoughts can differ. What to wear, where to shoot, how to posture are some of family photo ideas….


I think Le-vel Thrive flaunts the two components that I look for whenever I review a product: a good lineup of ingredients and positive feedback from people who have already tried it. Sure its price tag is a bit higher compared to its counterparts, but I think this one deserves to be in your shortlist of supplements to check out.
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text:
HomeSportsFootballAgnelli Admits Super League Cannot Go Ahead After Nine Clubs Pull Out
SportsFootball
Agnelli Admits Super League Cannot Go Ahead After Nine Clubs Pull Out
(The Guardian) The European Super League founder and Juventus chairman, Andrea Agnelli, has said that the breakaway can no longer go ahead after nine clubs withdrew.
Asked whether the project could still happen after the exits, Agnelli told Reuters: "To be frank and honest no, evidently that is not the case."
Agnelli insisted he remained convinced that European football needed to change and had no regrets about the way the breakaway attempt was made. "I remain convinced of the beauty of that project," he said, stating it would have created the best competition in the world. "But admittedly … I mean, I don't think that that project is now still up and running."
The Italian claimed other clubs were ready to join the new European league before the English sides left. "I'm not going to say how many clubs contacted me in just 24 hours asking if they could join," he said, declining to name them. "Maybe they lied, but I was contacted by a number of teams asking what they could do to join."
Counselors in Nevis Better Equipped to Treat Young Persons Who Abuse Cannabis
In Annual Address, Putin Warns Russia's Foes Will Be Sorry


Do we just get together and run?
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Coping with the Monster severely examines city governance in Africa, with specific connection with the intense difficulties and demanding situations posed via waste administration. It describes, compares and appraises the events in Abidjan, Dar es Salaam, Ibadan and Johannesburg, characterizing general varieties of governance and their successes and screw ups in facing the severe challenge of mounting city waste.
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OEC - Climate and Infrastructure I: Why Does It Matter?
The US National Academy of Engineering (NAE) has collaborated with four partners -- Arizona State University, Museum of Science-Boston, University of Virginia, and the Colorado School of Mines -- to develop a Phase I Climate Change Education Partnership (CCEP) on the impacts of climate change for engineered systems. In its work from 2010 to 2014 the Partnership focused on efforts to catalyze and transform education so that it effectively prepares current and future engineers, policymakers, and the public to meet these challenges. At the capstone meeting and workshop in January 2013, national leaders in climate adaptation, city management, engineering systems, public engagement, and other key fields gathered to work through the implications of climate change for US infrastructure in the context of education and public policy.
The workshop was recorded, and two videos released by the NAE Center for Engineering Ethics and Society summarize the results. The first highlights the importance of changes in climate with respect to engineered systems, and the second describes the many different types of expertise that will be needed to adequately address the challenges to these systems. These videos are intended to serve as a starting point for discussion and enable educators and others to inform engineers and the public of the numerous challenges posed by climate change.
This work was supported by Grant No. 1043289 between the National Academy of Engineering and the National Science Foundation. Any opinions, findings, conclusions, or recommendations expressed in these videos are those of the speaker(s) and do not necessarily reflect the views of the organizations or agencies that provided support for the project.


text:
\section{Introduction}

Hadron-hadron scattering at high energy is dominated by the exchange of weakly interacting Pomerons and Reggeons with vacuum and meson
quantum numbers respectively. Many decades ago, a
description of this process within the effective Reggeon field theory was pioneered by Gribov and collaborators~\cite{Gribov:1973jg}.

A first principle approach to Reggeon physics in the context of weakly coupled
QCD was succesfully obtained by re-summing rapidity ordered Feynman graphs \cite{Kuraev:1977fs}.
The BFKL Pomeron~\cite{Kuraev:1977fs}   emerges in leading order through the re-summed two-gluon
ring diagrams with vacuum quantum numbers. Few non-perturbative approaches
to Reggeon physics in the context of QCD have also been attempted, ranging from
the stochastic vacuum~\cite{PIRNER} to a Reggeized graviton in holography~\cite{Brower:2006ea}.
Here, we will be  interested in  the Pomeron as a semi-classical  stringy
instanton or in short the BKYZ Pomeron~\cite{Basar:2012jb}.

The BKYZ Pomeron carries intrinsic temperature  and entropy~\cite{Shuryak:2013sra},
i.e. $T=\chi/(2\pi b) $ with $\chi={\rm ln}(s/s_0)$ being the relative rapidity of the pair of  hadrons scattering at the
impact parameter $b$. As a result, a new phenomenon in hadron-hadron scattering at large
energy may take place when the intrinsic temperature of the string becomes comparable to
the Hagedorn temperature. A highly excited and long string with large energy and entropy
becomes a string ball, which, when cut, can lead to the large multiplicity states
observed in hadron-hadron energies at collider energies.

In this paper we will focus on the systematics of the QCD effective string theory (EST),
when the hadron-hadron scattering at large $\sqrt{s}$ is dominated by exchange of weakly
fluctuating  strings in a tube configuration.
 In section \ref{sec_pot} we review what is known about QCD strings from two sources:
 (i) the  effective  string theory (EST) and (ii) numerical studies of lattice gauge theory.
Each is focused on the static potential between point-like  and heavy color charges.
In section \ref{sec_duality} we  use the string duality relation, and translate those results from the
static potential to the Pomeron scattering amplitude. The phase of the Pomeron is addressed both
empirically and theoretically in the context of the BKYZ Pomeron in section IV. The evidence of the
odderon from the time-like or $t>0$ region is reviewed in section V using the latest lattice results.
A charge odd analogue to the charge even Pomeron is identified as the stringy odderon.  In section
VI  and VII we discuss the empirical evidence for the size and shape of the Pomeron at collider energies,
and suggest their theoretical descrptions using the EST of the Pomeron. Our summay and
conclusions are in section VIII. In the Appendix we briefly detail the unitarization of the string
amplitudes for fixed signature  in the EST.

\section{Effective string theory and the static potential } \label{sec_pot}

At large distances, the leading contribution to the static and heavy quark-antiquark
potential $V_0(r)$ in pure Yang-Mills theory is the famous linear potential

\begin{equation}
\label{0}
V_0(r) =\sigma_T r
\end{equation}
with $\sigma_T$ as the fundamental string tension.
In QCD with light quarks
this behavior is only valid till some distance due to screening by light quarks
in the form of two heavy-light mesons. This effect will be ignored below.

Long strings are described uniquely by the Polyakov-Luscher action~\cite{Polyakov:1986cs,Luscher}
(an expanded form of the Nambu-Goto action) in the form
\begin{equation}
\label{1}
S=-\sigma_T \int_M d^2x \,(1+\partial_\alpha X^i \partial^\alpha X_i)
\end{equation}
The integration is over the world-volume  of the string $M$ with embedded coordinates $X^i$
in D-dimensions. The first contribution is the area of the world-sheet, and the second contribution
captures the fluctuations of the world-sheet in leading order in the derivatives.

Since the QCD string is extended and therefore not fundamental, its description in terms
of an action is  ``effective" in the generic sense, organized in increasing derivative contributions
each with new coefficients.  These contributions are generically split into bulk $M$ and boundary
$\partial M$ terms. The former add pairs of derivatives to the Polyakov-Luscher action. The first of such
such a contribution in the gauge fixed as in (\ref{1}), was proposed by Polyakov~\cite{Polyakov:1986cs}

\begin{equation}
\label{2}
+\frac 1\kappa \int d^2x\, \left(\dot{\dot {X}}^\mu\dot{\dot{X}}_\mu +2\dot{X}^{\prime \mu}\dot{X}^\prime_\mu+
X^{\prime\prime\mu}X^{\prime\prime}_\mu\right)
\end{equation}
which is seen to be conformal with
the dimensionless extrinsic curvature.
Higher derivative contributions are restricted by
Lorentz (rotational in Euclidean time) symmetry. The boundary contributions are also restricted
by symmetry. The leading contribution is a constant $\mu$, plus higher derivatives. We will only
consider the so-called $b_2$ contribution with specifically

\begin{equation}
\label{3}
S_b=\int_{\partial M} d^2x \left (\mu +b_2\, (\partial_0\partial_1 X^i)^2\right)
\end{equation}

All the terms in  (\ref{1}-\ref{3})  contribute to the static potential (\ref{0}). The first contribution stems from the string
vibrations as described in the quadratic term (\ref{1}),

\begin{equation}
\label{4}
\sigma_Tr\left(1+\frac {V_0}{\sigma_Tr^2}\right)
\end{equation}
It is Luscher universal term with $V_0=-\pi/12$ in 4-dimensions~\cite{Luscher}.  Using string dualities,
 Luscher and Weisz~\cite{Luscher:2004ib} have shown that the next and higher contribution is universal

\begin{equation}
\label{5}
\sigma_Tr\left(1+\frac {V_0}{\sigma_Tr^2}-\frac 18 \left(\frac{V_0}{\sigma_Tr^2}\right)^2\right)
\end{equation}
These contributions are part of a string of contributions re-summed by~Arvis \cite{Arvis:1983fp}
using the Nambu-Goto action

 \begin{equation}
\label{6}
V_{\rm Arvis }(r)=\sigma_T r \left(1-{\pi \over 6} {1 \over \sigma_T r^2}\right)^{\frac 12}
 \end{equation}
For further discussion of the static $Q\bar Q$ potential stemming from the EST we refer to recent work in~\cite{Aharony:2010db}.  To order $1/r^4$ all the bare contributions to the potential are known

\begin{equation}
\label{VB1}
V(r)\approx \sigma_T r -\mu-\frac {\pi D_\perp }{24 r}-\frac{\pi^2}{2\sigma r^3}\left(\frac{D_\perp}{24}\right)^2 +{\tilde b_2 \over r^4} +...
\end{equation}
with general number of transverse dimensions $D_\perp=D-2$.
The $\mu$ term receives both perturbative and non-perturbative  contributions.
The former are UV sensitive and in dimensional regularization renormalize to zero,
as we  assume throughout. The latter are not accounted for  in the conformal  Nambu-Goto string,
but arise from the extrinsic curvature term (\ref{3}) in the form~\cite{Hidaka:2009xh,Qian:2014jna,footnote}.

\begin{equation}
\label{VB2}
\frac{D_\perp}4\sqrt{\sigma\kappa}\rightarrow \mu
\end{equation}
Note that this contribution amounts to a negative boundary mass term in
(\ref{3}), and vanishes for $D=4$ spacetime-dimensions. It
 is finite for $D_\perp> 2$ in the holographic AdS/QCD approach. The third and fourth contributions in (\ref{VB1})
are Luscher and Luscher-Weisz universal terms in arbitrary dimensions, both reproduced by expanding
Arvis potential; see \cite{Petrov:2014jya} for a related discussion of the role of Luscher terms in the Pomeron structure. The last contribution is induced by the derivative-dependent string boundary contribution (\ref{3}).
(Note that even if the string ends are constant in
external space, they still may depend on the  2-dimensional  coordinates on the worldvolume of a vibrating string.)

We will not discuss the extensive holographic studies of the EST  and related  potential~\cite{Aharony:2010db},
but proceed to lattice simulations of the heavy-quark potential. These studies have now reached
a high degree of precision, shedding  light on the relevance and limitation of the string
description.  In a recent investigation by Brandt~\cite{Brandt:2017yzw} considerable accuracy was
obtained for the potential at zero temperature and for pure gauge SU(2) and SU(3) theories.
As can be seen from Fig.~3 in~\cite{Brandt:2017yzw}, the inter-quark potential is described to
an accuracy of one-per-mille, clearly showing that both  Luscher's universal terms, $1/r, 1/r^3$
are correctly reproduced  by the numerical simulations. Indeed, for  $r/r_0>1.5$ (or $r>0.75\,$  fm
for Sommer's parameter $r_0=0.5\,$ fm) these two contributions  describe the potential
extremely well.

Expanding further to order  $1/r^5$, or keeping the complete square root in  Arvis potential
(\ref{6}), would not improve the agreement with the lattice potential, since the measured
potential turns up and opposite to the expansion. Brandt  lattice simulations~\cite{Brandt:2017yzw}
have convincingly demonstrated that the next correction is  of order $1/r^4$ with the opposite sign.
The extracted contribution fixes the $b_2$ coefficient   in (\ref{VB1}) as

\begin{equation}
\tilde b_2= - {\pi^3 D_\perp \over 60 } b_2
 \end{equation}
with the numerically fitted values

\begin{eqnarray}
b_2^{SU(2)} \sigma_T^{3/2} = && -0.0257 (3)(38)(17)(3) \nonumber\\
b_2^{SU(3)} \sigma_T^{3/2}=&&-0.0187 (2)(13)(4) (2)
\end{eqnarray}
(for the details and explanation regarding  the procedure and meaning of the errors we refer to~\cite{Brandt:2017yzw}).
Note that the overall contribution of this term to the potential is positive.
 What this means  is that at $r\approx r_0=0.5\,$ fm the static potential
contains a wiggle, visible however only with a good magnifying glass since its relative magnitude is  $10^{-3}$. We will
now explain how this impacts the scattering amplitude of color singlet dipoles at large $\sqrt{s}$.

\section{From the static potential to the  Pomeron} \label{sec_duality}

The (Euclidean) world-volume of the string for a static and heavy inter-quark  potential is a rectangle with the size
$\hbar/T$ in time and $r$ in space. It is assumed that $r \ll \hbar/T$ and the string is not excited.
In contrast, the  BKYZ Pomeron ~\cite{Basar:2012jb} is derived from a stringy instanton
with the world-volume of a shape of a ``tube", with a mean circumference $\beta=2\pi b/\chi$
and length $b$, the impact parameter. In this case a string is little excited when
it is long in space, $b \gg \beta$, the opposite of the condition above.

As discussed in the Appendix of~\cite{Shuryak:2017phz}, one can
map the two problems at hand via some duality relation, by exchanging time and space, and also by
adding another mirror image of a potential and match the boundary conditions.
In this case the two partition functions of the string and its excitations become identical.
The explicit transformation is

\begin{equation}
2b \leftrightarrow {\hbar \over T}, \,\,\, \,\, \beta \leftrightarrow 2r
\end{equation}
Assuming the correspondence between the potential and the Pomeron is exact
we can map the potential (\ref{VB1}) onto
the Pomeron scattering  amplitude in in impact parameter space  as

\begin{equation}
\label{VB3}
{\cal A}{(\beta, b)}\approx 2is\,{\bf K}\approx 2is \,e^{-{\cal S}(\beta,b)} \end{equation}
with explicitly

\begin{eqnarray}
\label{SPOT}
{\cal S}=&&+\sigma\beta b-2\mu b-\frac{\pi D_\perp}{6}\frac b\beta\nonumber\\
&&-\frac{8\pi^2}{\sigma}\frac{b}{\beta^3}\left(\frac{D_\perp}{24}\right)^2
-{ 2^5 b \tilde b_2\over \beta^4}
\end{eqnarray}
with $\sigma=\sigma_T/2$ and $2\pi\sigma_T=1/\alpha^\prime_{\mathbb R}$.
Now, we can recall the parameters of the ``tube" and set
 $\beta=2\pi b/\chi$, with the shorthand notation

\begin{equation}
\chi={\rm ln}(s/s_0)
 \end{equation}

 Following this substitution, one observes that the leading and subleading  terms
have very different roles and  energy dependence. For instance, the leading two contributions
with a rapidity dependence in (\ref{VB3}) are

 \begin{equation}
 \label{VB4}
e^{\chi\frac{D_\perp}{12}-\frac{b^2}{4\chi\alpha_P^\prime}}
\end{equation}
with the  Pomeron intercept of $\Delta=\frac{D_\perp}{12}$,
and the term responsible for the Pomeron slope. The latter yields the famous Gribov diffusion,
at the origin of the large-distance growth of the hadrons according to the
``diffusive law"
\begin{equation}
b^2 \sim \chi ={\rm ln}\left({s\over s_0}\right)
\label{eqn_diff}
 \end{equation}
which exists equally for perturbative gluons and strings.

Furthermore, one should recognize that the stringy Pomeron approach exists in two versions, the flat space  and the holographic
ones. In the former case the space has two flat transverse directions $D_\perp=2$,
while in the latter the string  also propagates in the third  and curved dimension. Since Gribov diffusion
also takes place along this coordinate, identified with the ``scale" of the incoming dipoles,
the expressions we will use are a bit modified from the standard expressions.
One such effect, derived for the  the  BKYZ Pomeron in~\cite{STOFFERS}, is the modification of the Pomeron intercept
\begin{equation}
{D_\perp \over 12} \rightarrow {D_\perp \over 12}\left(1-{3(D_\perp -1)^2 \over 2D_\perp \sqrt{\lambda} }\right)
\label{eqn_Delta_P}
\end{equation}
 Here $D_\perp=3$ and $\lambda=g^2N_c$ is the 't Hooft coupling, assumed to be  large.
 In the range of $\lambda=20-40$,  (\ref{eqn_Delta_P})  is in the range 0.14-0.18. For the numerical analyses to
 follow, we will use for the Pomeron intercept the value $\alpha_{\mathbb P}(0)-1=\Delta_{\mathbb P}=0.18$.
 (This  happens to be  not far from the flat space value of $\frac 16=0.166$.)

At the end of this section, let us discuss the way we fix the absolute units used in this work.
There is a dilemma,  well known in the literature.  One textbook approach is to rely
on the slope of the mesonic/baryonic trajectories $\alpha^\prime_{\mathbb R}=1/(2 \pi \sigma_T)\approx 0.9 \, {\rm GeV}^{-2}$,
related with the fundamental tension of the open strings. If so, the closed strings glueballs
should have trajectories with $\alpha^\prime_{\rm glueballs}=\alpha^\prime_{\mathbb R}/2\approx 0.45 \, {\rm GeV}^{-2}$
as the tension is doubled. Using those canonical values is one option to fix the scale.
However, multiple fits to the Pomeron produce different and much smaller values for $\alphaÕ_{\mathbb P}$. The earlier fits
 in~\cite{Burq:1982ja} yield $\alpha^\prime_{\mathbb P}=0.14\pm 0.03\,{\rm GeV}^{-2}$, while the later ones
 in~\cite{Donnachie:1992ny} yield $\alpha^\prime_{\mathbb P}=0.25\,{\rm GeV}^{-2}$. We note that the string tension
is warped in holography, and one would expect an effective $\sigma_T$ and therefore an effective $\alpha_R^\prime$ on average. This issue will not be addressed here.

In our previous paper \cite{Shuryak:2013sra} we used the results of lattice gauge
simulations ~\cite{Meyer:2004gx}, as reproduced in Fig.\ref{fig_Cpositive}, which
explains this deviation by the fact that the leading trajectory -- apparently unlike
the second one -- is not linear but has a quadratic term. Our fitted value for $\alphaÕ_{\mathbb P}=0.20\, {\rm GeV}^{-2}$.
We will be using this value to fix the units in this work.

\begin{figure}[h!]
\begin{center}
\includegraphics[width=7cm]{Gluebals_Cpositive}
\caption{Positive C-parity glueball according to lattice gauge
simulations  \cite{Meyer:2004gx}. The lines are our fit to Regge trajectories
\cite{Shuryak:2013sra}. }
\label{fig_Cpositive}
\end{center}
\end{figure}

  \begin{figure}[htbp]
\begin{center}
\includegraphics[width=7cm]{angles}
\caption{Euclidean rendering of the $qq$ scattering amplitude at relative angle $\theta$ and the $q\bar q$
scattering amplitude at angle $\theta+\pi$. The continuation to Minkowski space follows through $\theta\rightarrow -i\chi$
with the latter as the charge conjugate of the former. The relevance of this construction for the Pomeron and the Odderon is detailed in the text.}
\label{fig_crossed}
\end{center}
\end{figure}

   \section{The phase of the Pomeron  }

Experimentally, the Pomeron scattering amplitude exibits both a real and imaginary part.
The real part is measured at two locations:  at  small $t\approx 0$, by observing the
interference with the weak Coulomb scattering, and   at the location  of
the diffractive peak where the imaginary part vanishes.  For the interference measurement,
the results are expressed in terms of the so called $\rho$ parameter
\begin{equation}
 \rho={{\rm Re}({\cal A}(s,t=0)) \over {\rm Im}({\cal A}(s,t=0)) }
\end{equation}
  The recent TOTEM data~\cite{Antchev:2016vpy} give
  \begin{equation}
\rho(\sqrt{s}=8\, {\rm TeV})=0.12\pm 0.03
\end{equation}
The textbook
description of the Regge  scattering amplitudes relates the $\rho-$parameter with the signature factor
which for small $t$ is captured by the phase factor $e^{-\frac i2\pi \Delta_{\mathbb P}}$. It is small
if $ \Delta_{\mathbb P}$ is small, in agreement with data.

We now note how similar terms appear in the stringy amplitudes, starting from the positive definite charge
signature $C=+1$, the Pomeron. Such a  phase appears when analytically continuing the Euclidean scattering amplitude
at angle $\theta$ to $\pi+\theta$ for the cross channel, as illustrated in Fig.~\ref{fig_crossed} for quark-quark  scattering.
The result after integrating over the impact parameter and keeping only the two leading
contributions in (\ref{SPOT}) is for the $C=+1$ signature

\begin{eqnarray}
\label{TPLUS}
&&{\cal A}^+(t,s)\approx 2is\left(\frac \pi 2g_s a_D\right)^2\nonumber\\
&&\left(\left({\rm ln}(s)\right)^{1-\frac{D_\perp}2}\,s^{\alpha_{\mathbb P}(t)} +
\left({\rm ln}(-s)\right)^{1-\frac{D_\perp}2}\,(-s)^{\alpha_{\mathbb P}(t)}\right)\nonumber\\
\end{eqnarray}
with $\alpha_{\mathbb P}(t)=\Delta+{\alpha^\prime_{\mathbb P}} t$. At large $\sqrt{s}$ and small $t$, (\ref{TPLUS}) is
also characterized by an extra phase $e^{-\frac i2\pi\Delta}$ in agreement with textbooks.
However, we note that while the pre-exponential factors with ${\rm ln}(s)$
disappear in 4-d flat space, i.e. $D_\perp=2$, as in the textbooks, they
do not for  $D_\perp> 2$ as is the case of the  holographic model
(see (\ref{eqn_Delta_P})). The additional pre-exponential factors
stem in this case from the diffusion in the additional holographic direction,
and are thus completely necessary. Their inclusion through the branch ${\rm ln}(-s)\equiv {\rm ln}(s)+i\pi$ produces
a correction to the phase
which is asymptotically  subleading at infinite $\chi$, but at current energies is about -15\%  (it enters
with the opposite sign to the standard phase.) So, accurate measurements of the energy dependence of the phase can be sensitive to Gribov diffusion in the 5-th dimension.

 \begin{figure}[h!]
\begin{center}
\includegraphics[width=6cm]{odderons2}
\caption{$C=-1$ parity glueballs on a Regge plot, showing
 their angular momentum $J$ versus their squared mass $M^2 ({\rm GeV})^2$.
  The two upper (blue) points and lines are for the negative spatial parity $P=-1$
  gluballs, and the lower (red) ones are for the $P=+1$.  The masses are from lattice gauge
simulations \cite{Meyer:2004gx}.
The lines shows the hypothetical behavior of the Regge trajectories.}
\label{fig_odd}
\end{center}
\end{figure}

  \section{ The Odderon}

Let us start with the phenomenology of glueballs.
One can get some information about the odderon Regge trajectory
following the same procedure as we used e.g. in~\cite{Shuryak:2013sra}.
Specifically, this consists in:  1/ plotting the  known masses (squared)
of the appropriate glueballs; 2/    connecting  those by plausible lines, to learn the
relevant slope of the trajectories; 3/  extrapolating the upper trajectory to $t=0$.
The corresponding plot for  the masses of  glueball with positive C parity, taken
from the lattice study by Meyer~\cite{Meyer:2004gx}, is shown in
Fig.~\ref{fig_Cpositive} for completeness.

It is straightforward to extend it to glueballs with
 negative $C$-parity  as well, as we show in Fig.~\ref{fig_odd}.
  First, we note that there are 4 glueballs with negative $P=-1$ (vector-like) and 4 with positive
  $P=+1$ (axial-like) spatial parity. As seen from this figure, it appears that
  6 of those 8 states make three pairs defining straight and parallel lines with the same (within errors)
  slope, namely $\alpha^\prime=0.32\, {\rm GeV}^2$. Furthermore, we note that the lowest (the leftest)
  vector $J^{PC}=1^{--}$ state seems to belong to the second trajectory. This feature is the same as for the positive $C$ parity case, in which the lowest
 scalar $0^{++}$ also belongs to the second trajectory. We take both features as
 indications that the trajectories are correct as drawn here.

 Now, with only two states on the upmost trajectory, one can only define uniquely a straight line.
 Extrapolation of those to $t=0$
 would indicate that the odderon intercept $\alpha_{\mathbb O}(0)$ is negative. If so, its contribution
 is  suppressed by at least one power of $s$ compared to the Pomeron, or
 $\sqrt{s}$ compared to the  mesonic Regge trajectories:
  in this case it is way too small to be observed at
 RHIC/Tevatron/LHC energies.

 However, we know from experiment
 (the scattering data at negative $t$ for the Pomeron) as well as from the
  positive $C$-parity plot (in which one knows 3 states, $J=2,4,6$)
  that the upper Pomeron trajectory is not linear but curved. One may speculate
  that the same feature would hold  for the odderon as illustrated by
the speculative dashed line in   Fig.~\ref{fig_odd}.

On the theoretical side, the stringy description of the Pomeron also allows for
a stringy odderon. The dipole-dipole scattering with negative signature follows
the same reasoning as that for the positive signature in (\ref{TPLUS}), with now the result

\begin{eqnarray}
\label{TMINUS}
&&{\cal A}^-(t,s)\approx 2is\left(\frac \pi 2g_s a_D\right)^2\nonumber\\
&&\left(\left({\rm ln}(s)\right)^{1-\frac{D_\perp}2}\,s^{\alpha_{\mathbb P}(t)} -
\left({\rm ln}(-s)\right)^{1-\frac{D_\perp}2}\,(-s)^{\alpha_{\mathbb P}(t)}\right)\nonumber\\
\end{eqnarray}
 At large $\sqrt{s}$ and small $t$, the amplitude (\ref{TMINUS}) is only parametrically
suppressed from (\ref{TPLUS}) by ${\rm tan}(\pi\Delta/2)$. Remarkably, for the critical
bosonic string with $\Delta=D_\perp/12=2$, the odderon amplitude vanishes identically.
This is not the case, for the holographic string with $2< D_\perp\leq 3$, where the suppression
is parametrically small (especially for our empirical value of $\Delta_{\mathbb P}=0.18$),
but with an odderon with the same intercept as the Pomeron.
The inclusion of the higher stringy corrections should not alter qualitatively these observations.

 Summarizing this section: the  negative $C$-parity glueballs hint towards the  existence of
 a set of Regge trajectories, with a slope different from any other set. If the
 leading trajectory is curved, the location of its intercept $\alpha_O(0)$
 is likely to lie between zero and one.  The odderon drops out from the critical
bosonic string, but otherwise carries a similar intercept as the Pomeron for the
non-critical and holographic string. While its relative contribution is  small, its relyable observation would also be an indication to diffusion in the 5-th dimension.

\begin{figure}
\begin{center}
\includegraphics[width=6cm]{cr_from_profile6e}
\includegraphics[width=6cm]{B_from_profile6e}
\includegraphics[width=6cm]{ratio_from_profile6e}
\caption{The upper plot shows $\sigma_{\rm tot}$ in $mb$ versus the log of the collision energy $\chi={\rm ln}(s/s_0)$.
The left-side (low energy) data points at $\sqrt{s}=27,63\, {\rm GeV}$  are from the old ISR measurements, and the three right-side points, for
$\sqrt{s}=2.76, 7, 8 \, {\rm TeV}$ are from the TOTEM measurements. The dotted line in the lower plot indicates the contribution of the Reggeons other than the Pomeron (from the PDG fit.)
The middle plot shows the elastic slope  $B$ (${\rm GeV}^{-2}$), and the lower plot
is their ratio. The curves are for the profile shown in Fig.~\ref{fig_profile_from_profile6}.
}
\label{fig_pheno_from_profile6}
\end{center}
\end{figure}

\section{The Pomeron shape  }

\subsection{Phenomenology}

The main information we have about the shape of the scattering amplitude $\Omega(s,b)$
can be summarized as follows. One observable is the total cross section, which
is related to it via the optical theorem
\begin{equation}
\sigma_{\rm tot}=4\pi \int db b \,{\rm  Im}\, \left( \Omega(s,b) \right)
\end{equation}
Some data points for the cross section are shown in
the upper plot of Fig.~\ref{fig_pheno_from_profile6}. As  well known, the cross section
more than double from ISR to LHC energies.
However, one should be aware that the low energy plots are affected by  the Reggeons other than the Pomeron.
Their contribution (from the PDG) is shown by the  the dotted line in the left corner.

Another important parameter is the elastic slope $B$
\begin{equation}
B(t=0)=\left(-\frac{d{\rm ln}\sigma_e}{d|t|}\right)_0=\frac 12 \left<b^2\right>_\Omega
\end{equation}
The corresponding data are shown  in
the middle plot of Fig.~\ref{fig_pheno_from_profile6}. Their value
also grows with the collision energy, which is due to the effective growth
of the proton size induced by Gribov diffusion process.

In order to understand what these data tell us about the shape of the
profile, it is useful also to plot their dimensionless ratio

\begin{equation}
\label{RATIO}
R\equiv {\sigma_{\rm tot}(s) \over B(s)}
\end{equation}
with the cross section expressed in ${\rm GeV}^{-2}$. This ratio is plotted in the lower plot in
Fig.~\ref{fig_pheno_from_profile6}.
In order to understand what the ratio (\ref{RATIO}) tells us, it is convenient to compare it to some simplified
models for the shape,  such as e.g. the {\em black disc}

\begin{equation} \Omega_{bd}=\theta(b_{max}-b) , \, \,\,R_{bd}=8\pi \end{equation}
the $Gaussian$ without a prefactor
\begin{equation} \Omega_{G}=e^{-b^2/2b_0^2}, \, \,\,R_{G}=4\pi \end{equation}
and the $exponential$, also without a prefactor
\begin{equation}  \Omega_{E}=e^{-b/b_0}, \, \,\,R_{E}={4\pi \over 3} \end{equation}

The high energy values of the ratio are far from the two extreme shapes, and rather close
to the Gaussian value. The low energy profiles are in fact also near-Gaussian,
but with the prefactor smaller than one, as we will discuss below.

\subsection{The Pomeron shape and the mass of the string's ends}
As we already noted above, the
shapes of the BFKL and BKYZ Pomeron scattering amplitudes $A(s,b)$ are both Gaussian, as  is
typical for diffusive processes.   The prefactor, however,  growing as $s^\Delta$,
at high enough  energy violates the unitarity.

A generic resolution for this situation is well known: one has
  to include the ``multi-Pomeron effect", see e.g.~\cite{Schegelsky:2011aa} and references therein.
 A standard procedure  is  Glauber unitarization, which
follows from the substitution
\begin{equation}
\label{GLAUBER}
 {\bf K} \rightarrow \Omega\equiv 1-e^{\bf -K}
\end{equation}
The two leading terms in ${\bf K}$ compensate each other when
\begin{equation}
 \chi \Delta \approx {b_{\rm black}^2 \over 4\chi \alpha^\prime_{\mathbb P}}
\end{equation}
For  $b< b_{\rm black}$ the interaction is too strong,
multiple Pomeron exchanges  screen each other, and the proton is effectively black.
Asymptotically, we have $b_{\rm black}\sim \chi^2$.

\begin{figure}
\begin{center}
\includegraphics[width=7cm]{profile_from_profile6e}
\caption{The  unitarized profile function $\Omega=1-e^{\bf -K}$, for $\sqrt{s}=8000 \, {\rm GeV}$ (upper blue curve)
and $62\, {\rm GeV}$ (lower red curve), versus the impact parameter $b$ (${\rm GeV}^{-1}$).
The solid curves at all plots are for the parameters $A=7.\,{\rm GeV}^{-2}$, $\mu =-0.04\, {\rm GeV}$, and the dashed ones are for
 $A=4.5\,{\rm GeV}^{-2}$, $\mu =0$.
 The dotted line in the lower plot indicate the contribution of the Reggeons other than the Pomeron.
}
\label{fig_profile_from_profile6}
\end{center}
\end{figure}

The purpose of this subsection is to incorporate these effects, and also see
whether the fit to the data may indicate a nonzero value of the
first higher order term $\mu$. A technical point is that
standard Reggeon expressions for the amplitude contains
the nucleon form factors $f_N^2(t)$ times the Regge factor $s^{\alpha(t)}$.
However, in order to perform unitarization (accounting for multi-Pomerons)
one has to proceed to the coordinate profile, and in this case the product involves the
convolution of functions,  which is unnecessary complicated.
A simple way out, sufficient for the purpose of this section, is to use
Gaussian form factors  $f_N^2(q)=e^{-A q^2}$
which allows us to derive  the coordinate profile analytically

\begin{eqnarray}
\label{KSB}
 {\bf K}(s,b)= C\,{\rm exp}\left(\Delta_{\mathbb P} \chi-{b^2 \over 4\left(A+\alpha^\prime_{\mathbb P} \chi\right)}+2\mu b\right)\nonumber\\
\end{eqnarray}
where the normalization constant $C=2.2$ is fitted to the TOTEM data, as all
 higher order stringy contributions decreasing with $b$ are for now neglected.
 We remark that while the occurence of $A$ in (\ref{KSB}) is exact for one-Pomeron
 exchange, it is only an approximation in the unitarized multi-Pomeron resummation
 in the Glauber substitution (\ref{GLAUBER}).

A negative $\mu$ shifts the maximum of the profile function away from $b=0$, and modifies the shape in a way shown
(for the unitarized profile functions)  in Fig.~\ref{fig_profile_from_profile6}.
It turned out that the shape of the amplitude is sensitive to even small nonzero $\mu$.

Now we invite the reader to look at Fig.~\ref{fig_pheno_from_profile6} again, now paying attention to the curves. In order to fit the upper plot, for the cross section, we used a smaller
Pomeron intercept $\Delta=0.09$. The middle plot, for the elastic slope, is however independent
of its value.
 One can see that the profile shape corresponding to $\mu=0$, the dashed line in the middle plot, describes the growth
 of  $B(\chi)$ quite well (but is not
good for the value of the ratio at the highest LHC point $\sqrt{s}=7\, TeV$ shown.
The solid line is better on that. The fact that it misses the low energy points
is less serious since one needs to subtract the non-Pomeron contributions.
 The upcoming LHC data for twice larger beam energy will clarify the situation, but for now
 we see that a small negative $\mu$ has a potential to make the description better.
   Perhaps the optimum,
 inside the simple ansatz for the shape we use, is somewhere in between these two curves.

  The purpose of these comparisons is not to get an excellent fit to the data, as there are
  already several of those on the market, using multi-parameter and rather arbitrary
  functions. Our main message here is that the profile shape -- and the two observables reflecting
 it --  are   very sensitive even to rather small $\mu$.  We
think this comparison tells us that its value is definitely small,  perhaps slightly negative.

There are also some theoretical justifications for a nonzero $\mu$, with a suggested sign.    In fact,  we already mentioned it in  (\ref{VB2}):
 the  extrinsic curvature term in the action does indeed produce such a contribution, see
~\cite{Hidaka:2009xh,Qian:2014jna}.
One may further ask if lattice studies of the potential can also help to find
the magnitude of $\mu$.
Unfortunately, on the lattice
 the constant  term in the potential $\mu\sim 1/a$ is singular in the $a\rightarrow 0$ continuum limit
 (a pointlike charge). A subtraction is possible, but the accuracy of the remaining finite part is
 not  good enough.

   Concluding this section, let us also mention another
   (admittedly much more exotic) origin for the   $B(s),\sigma(s)$ growth with energy:
    a hypothetical $repulsive$ string self-interaction, briefly discussed in \cite{Shuryak:2017phz}.

   \subsection{The  Pomeron profile and the Hagedorn transition}

In our paper~\cite{Shuryak:2013sra} we discussed the transition
between the stringy and perturbative regime, using the so called Hagedorn transition
in which strings get highly excited.
At the beginning of this paper we
have discussed the static potential, in which a specific structure  -- the ``wiggle" indicating a transition between
two regimes -- has been identified.
Furthermore, we argued that a stringy duality can connect the static potential
and the scattering amplitude, in a well defined way.

Now, let us discuss the issue phenomenologically. Suppose a ``wiggle"
is also present in the Pomeron profile function: what observables should be
used to locate it?

Our first comment is that $pp$ collisions at LHC energies already have a certain structure
in its unitarized profile function, namely the black disc. Obviously any
structure in the amplitude inside the black disc, for $r<r_{\rm bd}$, is unobservable.
To go around this difficulty one can either return to much lower energies of $pp$ collisions,
or focus on $\gamma p$ collisions at the future Electron-Ion Collider.

Suppose now that the ``wiggle"  happens at impact parameters $outside$ the black disc
region. For simplicity, let us imagine a small peak in the profile function at certain
value of the impact parameter $b_{\rm peak} > r_{\rm bd}$. A
Fourier-Bessel transform to momentum space will produce an oscillating signal
$\sim J_0(q b_{\rm peak})$. Its minimum is at $q b_{\rm peak}\approx 4$, and the
second peak is at $q b_{\rm peak}\approx 7$, not far from $2\pi$.

In fact TOTEM at 8 TeV does observe certain non-Gaussian deviations of the elastic peak,
parametrized by a simple Gaussian $e^{{\rm const}* t}$, with  a minimum located
at $q^2\approx 0.1 \, {\rm GeV}^2$. If this corresponds to a structure in the profile function,
its location should be at
\begin{equation} b_{\rm peak}={4 \over 0.316 \, {\rm GeV}}\approx 2.5 \, {\rm fm}\end{equation}
This distance is  way too far for the location of the Hagedorn transition.
The corresponding ``tube" temperature is way too low, and cannot correpond
to this transition.

  \section{Summary}

The effective string theory suggests corrections to the fundamental string Lagrangian, and elucidates how these corrections modify the linearly rising potetial at shorter distances. QCD lattice numerical simulations do  indeed
reveal   two universal ``Luscher terms"
  and a new non-universal term at even shorter distances
responsible for the appearance   of a  ``wiggle" in  the potential. Albeit  small in amplitude,  it
clearly marks
 the transition from the perturbative to the stringy regime.

The description of the static heavy quark potential in QCD can be mapped onto
the Pomeron scattering amplitude by string duality.
We have suggested that  the scattering amplitude can provide  a more direct
 test of the string dynamics than the static potential. We have discussed the effects predicted
 by those approaches, for the Pomeron size and shape.  The future experiments at the Electron-Ion Collider should allow,
 through the measurements of differential cross sections of $\gamma^* p$ scattering,
  to elucidate the structure of the Pomeron and
 more accurately define the domains of its pQCD and string descriptions.

  \vskip 1cm
{\bf Acknowledgements.} We thank V.A.Petrov for useful comments on the first version of this manuscript.
This work was supported by the U.S. Department of Energy under Contracts No. DE-FG-88ER40388 and DE-AC02-98CH10886.
\vskip 0.5cm

\section{Appendix: Unitarization for fixed signature}

The fully unitarized scattering amplitudes of given charge conjugation signature ${\cal T}^{C=\pm}$
in impact parameter space, follows from (\ref{VB3}) in terms of ${\bf K}$,  and its charge conjugate $\chi\rightarrow \chi-i\pi$
in terms of $\tilde{\bf K}$.  Specifically, for fixed impact parameter we have

\begin{eqnarray}
\label{APP1}
&&{\cal T}^++{\cal T}^-=2is\left(1-e^{-{\bf K}+\tilde{\bf K}}\right)\nonumber\\
&&{\cal T}^+-{\cal T}^-=2is\left(1-e^{-{\bf K}-\tilde{\bf K}}\right)
\end{eqnarray}
 or equivalently

\begin{eqnarray}
&&{\cal T}^+=+2is\left(1-e^{-{\bf K}}{\rm cosh}\,\tilde{\bf K}\right)\nonumber\\
&&{\cal T}^-=-2is\,e^{-{\bf K}}\,{\rm sinh}\,\tilde{\bf K}
\end{eqnarray}
The total cross sections for $pp$ and $p\bar p$ scattering can be deduced
from the forward scattering parts of (\ref{APP1})

\begin{eqnarray}
\label{APP2}
&&\frac 12 \left(\sigma_{p\bar p}+\sigma_{pp}\right)=\frac {4\pi}s \,{\rm Im}\,\tilde{\cal T}^+\nonumber\\
&&\frac 12 \left(\sigma_{p\bar p}-\sigma_{pp}\right)=\frac {4\pi}s \,{\rm Im}\,\tilde{\cal T}^-
\end{eqnarray}
where the Fourier transform at $t=0$  is subsumed in the amplitudes in (\ref{APP2}).


text:
Eurhythmics for Autism and Other Neurophysiologic Diagnoses
A Sensorimotor Music-Based Treatment Approach
Dorita S. Berger
Foreword By
Stephen M. Shore
In Eurhythmics for Autism and Other Neurophysiologic Diagnoses, Dorita S. Berger reveals how Eurhythmics, a method of teaching the musical concepts of rhythm, structure and expression kinaesthetically through movement, can help develop sensorimotor skills in children and adults with autism and other special needs.
Covering both theory and practice, she explains this innovative, music-based approach and how it can also address cognitive and sensory issues in adults with debilitating conditions, such as dementia or post-traumatic stress disorder. With a particular emphasis on autism, she provides clear and adaptable session plans, suitable for working with children and adults of all ages.
Press reviews for: Eurhythmics for Autism and Other Neurophysiologic Diagnoses
Stephen M. Shore, Clinical Assistant Professor of Special Education, Adelphi University
In a book like no other, Dr. Berger masterfully introduces the science behind the power of music and movement, followed by artistically presented, easy-to-implement, and practical intervention solutions for the clinician, parent, or other person to connect with individuals with autism and other neurophysiological diagnoses.
Frances H. Rauscher, Professor Emerita, University of Wisconsin Oshkosh, and co-author of 'Neurosciences in Music Pedagogy'
Dorita Berger's book is a gift to all individuals on the autism spectrum. As a music therapist and scientist, Dr. Berger combines cutting-edge research with a healthy dose of practical advice to show how eurhythmics in a music therapy-based clinical environment can be used to support a wide range of neurophysiologic diagnoses. This book is an outstanding resource for therapists, neuroscientists, and families.
Dr. Patricia Gray, Director of The BioMusic Program, University of North Carolina, Greensboro, USA
Dori Berger is that rare combination of musician, scientist and clinician. Based on her extensive research experience, she unfolds a detailed prescription of protocols, culled from Eurhythmics and the most recent science, that can lead to dramatically improved results. In doing so, she makes a clear argument for how music moves from a passive therapy to a potent intervention!
Colwyn Trevarthen, PhD, FRSE, Professor (Emeritus) of Child Psychology and Psychobiology, Department of Psychology, The University of Edinburgh, UK
A professional musician and practitioner in sensorimotor treatment of autism and related diagnoses, Dorita Berger writes brilliantly to inspire therapists who help persons with disorders of intention and awareness share life with self-confidence and joy. With up-to-date brain science of intentions and feelings, this book will also be a resource for students of psychology, medicine and education. Born musical, our sense of others' rhythms and melodies of affection, invention and discovery can be enhanced to overcome confusion and anxiety.
Bethan Habron-James
Approaches: An interdisciplinary journal of music therapy
[I]t was with great anticipation and enthusiasm that I welcomed the news of an addition to the literature in Dorita S. Berger's Eurhythmics for Autism and Other Neurophysiologic Diagnoses: A Sensorimotor Music-Based Treatment Approach... Reading this book as a specialist in Dalcroze Eurhythmics, I learnt a great deal about the science behind what I do... Her work focuses on finding a process of adaptation for those with special needs, which allows them to respond with courage to the demands of their environment, and to life in general... The scientific theories introduced in the opening section permeate the second section; the reader is led to understand why a particular exercise has a particular effect... As a whole, this is an accessible and inviting text to read... the reader is taken on a journey of understanding through carefully placed introductions and summaries... This is a book for learners, for those seeking to deepen their practise and who welcome the new. I encourage you to read it.


Rakhine state in Myanmar's west is riven with ethnic and religious divisions.
"We've been hearing gunshots and heavy artillery fire since about 7:30pm," local historian Khin Than told AFP by phone on Monday evening, adding that she saw soldiers shooting from about 10 vehicles as they drove around town.
"We dare not go outside and are hiding in our homes," she said. "We do not want our ancient treasures destroyed in the fighting."
"We are stuck in the middle," one Rohingya man told AFP, preferring not to give his name.
"Our injured villagers cannot go anywhere for treatment because of the situation. We can't do anything," he said.
The rebels enjoy widespread support from many ethnic Rakhine, who have felt marginalised for decades in one of the country's poorest states.
It is feared the expected verdict on Tuesday in the trial of a popular Rakhine politician for alleged treason could be a further flashpoint in Sittwe.


text:
from enum import Enum

class Color(Enum):
    beige = 1
    white = 2
    red = 3
    dark_orange = 4
    green = 5
    yellow = 6

    @staticmethod
    def from_string(value):
        if value is None or value == "":
            return None
        if value == 'beige':
            return Color.beige
        if value == 'white':
            return Color.white
        if value == 'red':
            return Color.red
        if value == 'dark_orange':
            return Color.dark_orange
        if value == 'green':
            return Color.green
        if value == 'yellow':
            return Color.yellow
        return None

redpajama_set_name: RedPajamaGithub