text:
Why Are School Libraries Essential?
Eight school librarians share their answers.
(What we do is more than offer flashy technology)
Your donations help support school libraries across the country.
Every child needs a safe place to fall — a place where they can explore things without worrying about failure and judgment. A library is one of those places. You can learn by following your nose in a library, which is very different from someone telling you what you should remember. Once a kid knows a library is hers to use as she wants, the world opens up.; I've seen it happen. It happened to me. -Bill Harley, Grammy-winning Singer, Songwriter, and Storyteller
School Librarian Day, National School Library Week, School Library Month; it is easy to get caught up with labels, titles, and spontaneous celebrations. However, do the above recognition days help define a school librarian's roles every day? It seems very easy to overlook the responsibilities of a school librarian. Luckily, first-hand accounts from school librarians in a comprehensive, geographic area help bring tasks into focus, which we have been able to use to represent the importance of school libraries for you.
Like sorting through countless holidays and celebrations that appear daily, weekly, and monthly, school librarians help others sort, organize, orient themselves, and create information.
Why are we telling you this? Here's how librarians can take a diverse range of needs and provide solutions in various ways based on their experiences, as told by school librarians around the country. (What we do is more than offer flashy technology).
School Librarians Spark Creative Sustainability
We just completed an outreach program with our public library partnership. The librarians came in, and our students learned a bit about hedgehogs. We crafted hedgehogs made out of weeded books. It was a great experience and allowed our students to do an activity that was visual, auditory, and kinesthetic! -Cindy, North Carolina
School Librarians Plan and Report Data
I think during Covid, we were finally recognized as having some expertise. Once the curriculum people realize we can share resources, they would invite us to the planning meetings. We also promote our programs and try and share positive things happening in the district and build newsletters. -CK, Wisconsin
School Librarians Oversee Curriculum Collation and Sources Interpretation
One of the newest collaborative tasks I have worked on recently has been the Library of Congress Grant given to the social studies team district-wide. I was asked to help collate sources from the Library of Congress for curriculum-based lessons. I don't think the administration and teachers know our expertise in this area. I was able to find sources more efficiently because I understood identifiers in the catalog that teachers are not trained to use. I collated forty-eight pages of modern history sources available digitally. In today's digital world, it is assumed everything is available digitally. As librarians, we can educate our staff and students on all sources. We can help them navigate and differentiate between credible and inaccurate information. -Kim, Delaware
Sign the pledge to vote for libraries!
School Librarians Foster Safe Spaces and Mentorship
Until recently, I have seen my role in high school as that of a mentor and provider of safe spaces. However, I have shifted my focus to change how our school and community view reading this year. Our test scores were not good, and the students needed to embrace reading. I plan to be the one to make this school one with a culture that reads. I have five years until retirement, and I have made drastic changes in the [library] environment to be welcoming. In hindsight, I realize that the environmental change and the relationships HAD to come first so that the students would trust me. Now, I have students who aren't "readers" coming to give it a try again. They say they haven't read since elementary school and are an accelerated reader. -Michelle, Georgia
School Librarians Are Community Connectors and Cultural Guides
My big goal has been to bring the library to life this year. I don't want students to think it's only a place for books. To that end, I've brought in interesting Chicagoans to bring our community and our city to the forefront at school. We have had visits from local authors, a snake expert, an architect, and a member of Blue Man Group. It gives kids a chance to see people living out the activities they read about in books and teaches them how to ask good questions and follow their curiosity. -Laura, Illinois
School Librarians Are Experts in Technology Liaising and Feng Shui Design
I have enhanced students' ability to do research by teaching them about databases, how to critique what they find on the internet and discern if it's true or not, and to help them vet sites on the internet to determine whether the information is reliable or not. I believe I am encouraging a love of reading (re-introducing them to books). At the high school level, they are so involved in other activities, like social media, sports, theatre, music, work, driving, etc. They let reading go. I introduced Sora (a student reading application of the Ebook provider, Overdrive) to school before the state started sponsoring it. My students could access books online via Sora and Overdrive during the pandemic and reached out to me for how to use them. I run book trailers on my TV in the library; I use social media to promote new books and my presence in the school. TikTok (a popular video-sharing application) is a favorite, and I think students view me differently because of it. My space is inviting and not restrictive. Students feel safe here. We have a single stall inclusive bathroom for students and a personal care cabinet with hygiene items for all students but primarily for our homeless and struggling population. Staff feels included in the library too. We have a coffee bar, kitchen, and copier here. -Debbie, Delaware
Take action today to support libraries!
How Does This All Connect?
Visually, how does this all connect and become represented visually? Check out Run Librarian, Run! from Nikki, a librarian from New Zealand who does just that. A former elementary librarian colleague with myself at the American School Foundation in Mexico City and supporting the InterBaccalaureate Program. Visually, Ms. Jayne captures the role a librarian fulfills daily outside the confines of what many do not realize about the librarian occupation. Now employed in New Jersey, Nikki finds there are no geographic borders to the skills school librarians perform daily.
Returning to the question, "what in the world would we do without our libraries?" so eloquently posed by actress Audrey Hepburn, we discover that based on the above insights, the answer to Ms. Hepburn's inquiry resides in each of us. Answering and magnifying answers to this pertinent question are possible during May's National Library Month. The responsibility of receiving and providing opportunities is up to us as lifelong learners and school librarians who define themselves as purveyors of technology, sifters of accurate information, and locksmiths of accredited professional communication. All you have to do is open the door.
Sign up to support school libraries across the United States.


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Young Odessa
March 30, 2016 Movie Reviews, Uncategorizedstevestevelovellidau
Talk about a mutual admiration society. Writing in the Oz, venerable reviewer/'The Movie Show' host David Stratton opined, 'I don't usually like to make predictions…But for once I'll stick my neck out…Odessa Young will be an international star.' On radio, during an interview for the ABC, the tender-yeared Ms Young told of her time at the Venice and Toronto film festivals last year. She reflected on the fact that she met and chatted with some of the world's best known film celebrities, but when Stratton approached her she became tongue-tied in awe. Here, in real life, was the man she grew up with during her fixation on hearing his opinion on the latest of moving-picture releases.
Attending two film festivals with her first two movies – what an amazing experience for a mere seventeen year-old. But the actress is no ingenue when it comes to acting. She has had a long apprenticeship in the television industry, most notably as a lead in Auntie's adaptation, for teens, of 'My Place'.
The two local products being showcased at the aforementioned exotic locales were both competently made – one considerably more worthy of a film festival than the other, in this humble scribe's view.
The lesser offering was 'Looking for Grace', the better 'Daughter'. In both Odessa Y could be said to have upstaged more seasoned old hands, such as Richard Roxburgh, Radha Mitchell and Terry Norris in the former. In the latter there was an even more stellar cast including Geoffrey Rush, Sam Neill and Miranda Otto. But whether she'll be our next Nicole Kidman/Cate Blanchett remains to be seen. But she's certainly off to a flyer.
Sadly, neither of her newly released vehicles seem to be setting the world alight at the box office, although 'Daughter' is still in the cinemas and may pick up through word of mouth. It would be a pity for it to not get the bums on seats it deserves.
'Looking for Grace' came to us from Sue Brooks, who presented with the marvellous 'Japanese Story' back in '03 – still one of my favs of the local product. Paul Byrnes, Age film critic, awarded 'Looking for Grace' a high rating of 4.5 stars, claiming it was as good, if not better, than 'JS'. Wrong Paul. 'Looking for Grace' isn't within a bull's roar. Saying that, it was watchable and did have its moments of pleasure – Roxburgh doing his almost, by now, compulsory rumpled/addled shtick, as well as when Norris was on screen. But it lacked the magic ingredient of 'Japanese Story' that audiences responded to. It lacked heart. Still, it was Young's debut and if Stratton is correct about her, that may be enough for it to be remembered and revisited.
'Daughter', on the other hand, is a different beast entirely. Novice director Simon Stone, as the great David S also predicts, is destined for greater glory based on the quality of this offering. It's very much an Aussie take on Scandi-noir – chilly landscapes and life-battered characters. And it's based on Norwegian playwright Ibsen's 'Wild Duck' – somewhat loosely. The drama is taught, tight and bleak. In a high country timber town, populated by brittle denizens, the local saw mill closing down impacts majorly. As well, some are carrying deep, dark secrets that come to the fore when wretchedly self-centred and alcoholic Christian (Paul Schneider) returns to Oz to attend the wedding of the town squire (Rush). His bride is a much younger woman, his former housekeeper in fact. It's an engaging turn from Anna Torv. His arrival sets in motion a series of revelations that will tear lives asunder. The climax is almost unwatchable; the ending not at all happy-ever-afters, nor are strands tied up neatly in a bow. This, though, unlike 'Looking for Grace', will stay with the viewer.
So David Stratton and I will watch Ms Young's future career with much interest. I suspect he's right. Obviously he's made a mark by being mostly right about such matters. And she is very at ease on the big screen, with her performances, particularly in 'Daughter', very brave. Worth seeing the movie for that alone. And she is still so very, well, young. If 'Daughter' does go the way of 'Looking for Grace' maybe Hollywood will come calling and will place her in something to make the wider world sit up and take notice. Odessa Young is the future.
Trailer -' Looking for Grace' = https://www.youtube.com/watch?v=7KboDXLZM3o
Trailer – 'Daughter' =https://www.youtube.com/watch?v=TaC-SrFdRZg
← Pug in the Landscape	The Legacy – Kirsten Tranter →

redpajama_set_name: RedPajamaCommonCrawl


text:
Carolina Eyck, född 26 december 1987 i närheten av Berlin, är en av få kvinnliga thereminvirtouoser.

Vid fem års ålder fick Carolina Eyck sin första klaviaturundervisning och ett år senare inledde hon sin violinutbildning. Hennes lärare på en specialskola för musik i Berlin var professor Ulla Scholz. Då Carolina Eyck senare bytte instrument till altfiol studerade hon istället under professor Ditte Leser.

Hennes komposition Physikalische Formeln vann 2005 en tävling för unga kompositörer. Verket Sciciani – Am wendischen Burgwall, för dragspel och stråkar, vann året därpå den offentliga radiostationen RBBs kompositörspris.

I augusti 2007 mottog hon ett musikstipendium för fiol från Kungliga Musikhögskolan i Stockholm.

Carolina Eyck har studerat theremin under Lydia Kavina och spelar på en Big Briar Theremin series 91A och på en Etherwave proffs-Theremin tillverkad av Robert Moog. Hon skrev även den första tyskspråkiga thereminkursen Die Kunst des Thereminspiels. Som thereminsolist har hon framträtt på flera musikfestivaler i Basel, Davos, Łódź och Budapest.

Litteratur
 Carolina Eyck: Die Kunst des Thereminspiels. SERVI Verlag, Berlin 2006, , EAN 4025 1187 0631 – tyskspråkig utgåva.
  Carolina Eyck: The Art of Playing the Theremin. SERVI Verlag, Berlin 2006,  – engelskspråkig utgåva.

Externa länkar

 Personlig webbplats

Tyska musiker
Kvinnor
Födda 1987
Levande personer

redpajama_set_name: RedPajamaWikipedia


The League of American Bicyclists is promoting Bike-to-Work Week from May 13-17 and Bike-to-Work Day on Friday, May 17, 2019.
National Bike Month includes an ever-expanding diversity of events in communities nationwide — but the biggest day of the month is Bike to Work Day.
Q: What is the best route for me to bike to work?
Biking to work takes planning. One of the most important parts of this planning is determining the safest route to ride.
Start by thinking about the route as you drive to work. If the road to work has viable shoulders then it can, dependent upon other factors, possibly be at least one part of your route.
After coming up with a tentative route, try driving that route during the time you might be on the road.
Look for possible danger points such as busy intersections, malls and strip malls with multiple entrance and exits, coffee & donut stops (can be very dangerous during commute time).
Be observant of the motor traffic clearance, is there a shoulder available?
Check the road surface, are there many grates, potholes or debris along the route?
Talk to others who commute to work on bicycles as they will already know some good routes.
Refer to the Transportation Alternative's map for possible routes.
Consider going out of your way a bit to get a safer route; the highway is straight but not always the safest.
Ride the route on Saturday or Sunday as a test run. The traffic will certainly not be the same but it will give you time to observe other factors. If you're planning to commute in the dark (with proper lights on the bike) try driving in the dark and look for well lit and quiet roads.
After you've developed your route make sure to let family, friends and co-workers know about it. You may need a pickup some time.
Q: Is riding during rush hour dangerous then normal?
Maybe, due to traffic volumn; But the danger can be minimized.
The primary way to reduce the danger of rush hour traffic is choosing a safe route which avoids as much traffic as possible.
Secondarily, if possible, shift working hours to avoid as much of rush hour as possible. Biking early in the morning on local streets, even in the dark (with proper lighting) is safer then rush hour. In the early morning cars are scarce and easily visible.
Also, when riding with others you will be more visible then riding alone; Having a group to ride to work is more enjoyable and safe at the same time.
If you have to ride during rush hour you may be surpised to find that the trip may not, due to the volume of cars on the road, be that much longer then if you drove.
Q: What if I work too far from home to ride my bike to work?
Using your bike to commute to work does not mean that you have to cycle all the way to work. Some bike commuters opt to drive their car to a Park & Ride lot, the mall or one of the local strip shopping centers. Not only will this make your bike commuting a little easier but can also cut out the most dangerous section of the commute. One lesson regarding this is not to park at the spots closest to the stores; merchants do not appreciate their closest parking spots being used all day by non-business related vehicles.
Some commuters drive their car to work with their bike on the rack and cycle home and back the next day. This also gives them the options of both having their car available at work and also bring all the clothing they need for work.
Q: What type of bike is best to use to commute?
Any bike will do! A Road bikes may be more efficient on the road then a Mountain Bike and that may be a factor if distance is great enough; A hybrid may be a good choice for your commute; remember, key to cycling is comfort and some folks will find the added stability of wider tires more comfortable on the road. If you choose to use a Mountain Bike consider having slicks (or less knobby tires) mounted to make the commute easier. Also the tires on a Mountain Bike may be a bit more resistant to the hazards found on the shoulder of the road.
Q: Do I need lights on my bike when I bike to work?
If you're planning to commute anytime near dusk its a good idea to have a set of lights. Even when its not so dark that you need a light to see the road you should have blinking lights both front (white or amber) and rear (amber or red). A good part of being safe is insuring that you are seen while on the bike. Rear lights will work best when mounted at center or on the traffic side of the bike. Front lights should be bright enough to prevent over-running the beam. A fast rider needs at least 15 watts of light for good illumination while a slower rider may do okay with less ... but you can never have too much. You want to be seen by the driver who might be thinking of turning in front of you from across the road. Along with active lighting there are other devices such a reflective tapes and straps that can catch the eye of drivers. Relective clothing is also an option for helmet covers, jackets and leggings.
Q: What equipment do I need to carry for bike commuting?
Cycling tools don't vary between commuting and general cycling; the cyclist should always try to be self-sufficient and carry a pump, a spare tube, and some basic tools. Using a water bottle may not be necessary dependant upon the distance you ride. Having a cell phone is a good idea; talking on it while cycling is not.
Q: How do I dress for riding and working? And how can I clean myself up when I arrive to work?
A helmet should worn regardless of the destination of any ride. Other cycling apparel may not be necessary for commuting dependant upon the distance and other comfort factors.
If you plan to ride in foul or cold weather then the appropriate clothing should be considered. Watching the forecast for the commute home, before leaving in the morning, is a good idea.
Some commuters carry their work related clothing with them on a daily basis using a knapsack, panniers or bike rack; if you have a small storage space at work then stocking up a few days clothing before riding makes the daily cycling commute more comfortable. Some clothing items such as shoes and belts can be left at work.
Some cycling commuters are lucky enough to have showers and changing areas at work, if you don't have the facilities and work up a sweat then keep a container of wash and drys in your office. A quick trip to the bathroom before starting work will be appreciated by both you and your co-workers.


Have you thought about writing a book? Do you just not know where to begin? Do you get writer's block just thinking about writing a book?
Best-selling nonfiction author David Kadavy shares his simple process for writing a book. Build confidence, ditch your inner critic, and finally write your book with simple habits you can start today.
You can listen to this short read in under an hour, so it won't get in the way of the one thing standing between you and your book: action!
Download today and make the book you've dreamed of a reality.
Comes with a free sample chapter of David Kadavy's latest audiobook, The Heart to Start.
©2018 Kadavy, Inc. (P)2018 Kadavy, Inc.
I want to start writing my next book now!
After listening to the audio version of the book, a new idea came to me for my next book and I have an urge to start writing it. The steps David Kadavy lays out in this book are so simple. If you know how to read, you'll be able to finish your first book in a month! Highly recommended!
Has some useful tips for practical writing habits. Worth a read although some parts might not be relevant, depending on what you are focusing on.


text:
What your are sharing
Bowling Green Pipe & Cigar's Spring Trifecta of Bourbon & Cigars. Who will be our Win, Place, & Show?
by Adminon March 3, 2017
By Admin, BuyLocalBG.com, BuyLocalBg@gmail.com/
Friday, March 3rd, 2017 6:00 AM CST
Bowling Green Pipe & Cigar's Spring Trifecta. Who will Win, Place, & Show?
Colonel E.H. Taylor Jr. Straight Rye Whiskey 100 Proof: Colonel Taylor is widely considered one of the founding fathers of the bourbon industry, fighting for the Bottled-in-Bond Act of 1897, nearly three decades after he purchased the now Buffalo Trace Distillery. Now he's honored with this Single Barrel Bourbon. While most rye whiskeys are made with corn as an ingredient, this is a unique rye whiskey made only with rye and malted barley. This spicy and full-bodied 100 proof Straight Rye Whiskey pays tribute to the classic American whiskey style and the Bottled in Bond Act of 1897. On the nose, this whiskey is full of dried fruit, black pepper and a touch of fresh dill. A small sip brings an array of flavors both sweet and savory with a terrific balance of dark spices and subtle caramel overtones. The finish is especially pleasing with an oaky dryness that lingers just long enough.
Padron Family Reserve "The Little Hammer" 50 Years. In Celebration of Padrons 50th Anniversar, the limited production Family Reserve 50 Years box pressed cigar is Buy Local Bowling Green's favorite cigar. Aged tobacco a minimum of 10 years and is available in both Natural & Maduro sun-grown wrapper.
Bulleit® Bourbon 10 Year Old: Bulleit® Bourbon 10 Year Old is Tom Bulleit's selected reserve. Aged in charred American white oak, select Bulleit® Bourbon barrels were set aside to age for 10 years. The result is a special expression of Bulleit that provides a rich, deep, incredibly smooth sipping experience.
Camacho American Barrel Aged Robusto tubos: Utilizing tobaccos aged in bourbon barrels for 5 months in the Dominican Republic, Camacho breaks the mold with their latest installment. With the spirit of America and the essence of a true frontier, Camacho American Barrel-Aged is truly a monumental cigar. It unites the hardworking spirit of Kentucky's bourbon maker's and Camacho's master builders. At its core is six-year-old Corojo leaves, American broadleaf, and Pennsylvania maduro that has been locked away in charred Kentucky bourbon barrels. In order to extract the characteristics of Kentucky bourbon and American Oak, Camacho's master builders take on this arduous process for five months straight. Wrapped in rugged, all-American broadleaf tobaccos, Camacho American Barrel-Aged delivers a richness and complexity accentuated by notes of oak and sweet bourbon aromas. This intense and fiery, medium to full-bodied blend is an unforgettable experience that blazes new frontiers in complexity.
Jefferson's Reserve – This is Jefferson's flagship bourbon where they define the Jefferson's taste: big and bold with hints of spice, toffee and leather. The Reserve is the perfect marriage of four different recipes of bourbon, one of which is an elder statesman. The result is bourbon that, like Thomas Jefferson, is sophisticated, complex and unforgettable.
AVO Classic no. 2- This cigar is the harmony of two cigar legends – Avo Uvezian and Master Blender Hendrik Kelner. Their collaboration offers an unforgettable experience of balance, enjoyment and refined taste. This inviting blend is composed using rare tobaccos that have been aged for 25 years.
Gift Cards are available which is another great reason to choose the local shop over online. Check out Bowling Green Pipe and Cigar onfacebook. Located at 434 East Main on fountain square you can reach them at 904-2285. Open Monday through Saturday 11:00 am-8:00 pm.
Tags: & Show?, Bowling Green Pipe & Cigar's Spring Trifecta of Bourbon & Cigars. Who will be our Win, Place
Book The BG Cheese Wagon For Your Next Event!
Five Great 4th of July Looks from Steel Mill & Co.
Welcome Our Newest Partners, The 30 Bird. Bowling Green's Local Whole30 Approved® meal service and caterer.
Story Page


Hiya. Thanks for reading the very first baby post here on le new blog!
...is about being home...or rather...it's about being content. It's about enjoying who, what, and where I am. I have lots of dreams and goals I've yet to see accomplished, but I don't want to miss life because I'm so focused on tomorrow that I zone out and miss the good things happening today.
A little while back as I was getting off the elevator at work, I said to some random person, "It's almost Friday!" And instead of the usual, "Thank goodness!" response, this wonderful person told me, "My dad always said not to wish your life away."
Even on the hard days I try and remind myself of this. No, I'm not perfect at it, but this (today, right now, wherever, whatever, whoever, however) IS life. I can choose to enjoy it or be a grump and live in a future that may or may not turn out the way I hope.
That's what I hope to share here. We'll see how it goes.


text:
package text

import (
	"github.com/golang/freetype/truetype"
	"golang.org/x/image/font"
	"golang.org/x/image/font/basicfont"
)

type Face interface {
	font.Face

	Data() FaceData

	Clone(data FaceData) Face
}

type FaceData struct {
	Size    float64
	DPI     float64
	Hinting font.Hinting
	Stretch font.Stretch
	Style   font.Style
	Weight  font.Weight
}

type ttfFace struct {
	font.Face
	font *truetype.Font
	data FaceData
}

func (self ttfFace) Data() FaceData {
	return self.data
}

func (self ttfFace) Clone(data FaceData) Face {
	return TTF(self.font, data)
}

func TTF(ttf *truetype.Font, data FaceData) Face {
	return ttfFace{
		Face: truetype.NewFace(ttf, &truetype.Options{
			Size:    data.Size,
			DPI:     data.DPI,
			Hinting: data.Hinting,
		}),
		font: ttf,
		data: data,
	}
}

type basicFace struct {
	font.Face
}

func (self basicFace) Data() FaceData {
	return FaceData{Size: 12, DPI: 72}
}

func (self basicFace) Clone(data FaceData) Face {
	return basicFace{self.Face}
}

func BasicFont() Face {
	return basicFace{basicfont.Face7x13}
}

redpajama_set_name: RedPajamaGithub


text:
Home / All works / Strings / Capriccio
score (PDF) (+€12.50)
Capriccio quantity
Capriccio for violin solo was composed in 1999, commissioned by the city of Rotterdam, as part of a project involving 24 Dutch composers who each wrote one capriccio which were then recorded on CD by 3 eminent violinists.
The violin is overwhelming in its musical and technical possibilities, which can be intimidating for a composer. Capriccio was written in an intuitive way, and consists of grooves that are stylistically related to Irish fiddler music and Bayou country blues, as well as more traditional classical styles. Capriccio was recorded by Benjamin Schmid and part of the 2CD album JacobTV – Suites of Lux – Basta Records.


text:
\section{Introduction}
Supersymmetric Yang-Mills (SYM) theories in higher-dimensional
spacetime are quite attractive from both theoretical and
phenomenological viewpoints. Such theories in ten-dimensional
(10D) spacetime arise as low energy effective theories of
superstrings, which are considered as the most promising candidates
for an ultimate unification theory of elementary particles
including gravitational interactions.
It is indicated that SYM theories in various spacetime
dimensions appear from D-branes in superstring theories.
Various phenomenological superstring models beyond the standard
model of elementary particles have been studied so far
(for a review, see \cite{Ibanez}),
many of which are based on SYM effective theories.
Even without mentioning superstring theories, SYM theories in
higher-dimensional spacetime themselves are interesting enough
for phenomenological model buildings beyond the standard model.

The standard model in four-dimensional (4D) spacetime may be
realized at a low energy by compactifying the extra dimensions.
In such a case, how to break higher-dimensional supersymmetry
and to realize a chiral spectrum is the central issue.
Toroidal compactifications with certain orbifold projections
are mostly studied. Orbifolds may be considered as some singular
limit of Calabi-Yau (CY) manifolds where most field contents,
their wavefunctions and couplings in 4D effective theories
are determined by geometric data. However, in general, such
data are complicated and it is difficult to determine the
CY metric explicitly, that makes phenomenological analyses
on CY manifolds qualitative but not quantitative.

Chiral spectra can be obtained not only on such nontrivial
geometric backgrounds, but also even on a simple toroidal
background with some gauge fluxes in Yang-Mills (YM) sector.
(see Ref.~\cite{Cremades:2004wa} and references therein).\footnote{
 Other geometrical backgrounds with magnetic fluxes were also
studied (see e.g. \cite{Conlon:2008qi}).}
In the latter case, chiral field contents and couplings are
determined by the fluxes. For example, Yukawa couplings
are determined by overlap integrals of the matter wavefunctions
whose analytic forms are determined as functions of
fluxes~\cite{Cremades:2004wa}. It is remarkable that the
numbers of generations for chiral matter fields are determined
by the fluxes, which provides a possibility that the flavor
structure of the standard model is essentially determined
by the YM-fluxes with sub-leading corrections from the
background geometry. The observed flavor structures of
quark and lepton masses and mixings would be realized by
wavefunction localizations in extra dimensions due to the
presence of magnetic fluxes~\cite{Cremades:2004wa,Abe:2008fi,Choi:2009pv}
yielding some discrete flavor symmetries~\cite{Abe:2009vi}.\footnote{
Similar non-Abelian discrete flavor symmetries are obtained
in heterotic orbifold models \cite{Kobayashi:2004ya}.}
It is also remarkable that a certain class of magnetized
D-branes can be regarded as the T-dual of some intersecting
D-branes~\cite{Cremades:2004wa,Blumenhagen:2000wh}.

Here we focus on such YM-flux backgrounds which preserve
(at least) 4D ${\cal N}=1$ supersymmetry out of the full
higher-dimensional supersymmetry such as ${\cal N}=4$.
In this case, field fluctuations around the background form
${\cal N}=1$ supermultiplets and their action is written by
superfields in ${\cal N}=1$ superspace, as indicated in
Refs.~\cite{Marcus:1983wb,ArkaniHamed:2001tb} in the case
without fluxes. In this paper we present a superfield
description of SYM theories in higher-dimensional (especially
10D) spacetime with magnetic fluxes in extra dimensions
preserving 4D ${\cal N}=1$ supersymmetry, and derive 4D
effective action for zero-modes in the superspace.
Such a superfield description allows us to construct
phenomenological models systematically and to analyze
detailed low-energy properties of them, such as particle
and superparticle flavor structures, features of Higgs
particles and dark matter candidates, which are required
for verifying these models by recent and upcoming data
obtained from high-energy experiments as well as
cosmological observations.

The sections are organized as follows.
In Sec.~\ref{sec:10d}, we rewrite the 10D SYM action in 4D
${\cal N}=1$ superspace in the case with magnetic fluxes in
extra dimensions, and derive zero-mode equations for superfields.
In Sec.~\ref{sec:4d}, the 4D effective action for zero-modes
are shown after a dimensional reduction. The action is extended
to the case with local supersymmetry (i.e. supergravity)
in Sec.~\ref{sec:sugra} where its moduli dependence is
also identified. A direction of model building based on
the effective action is indicated in Sec.~\ref{sec:model}.
Finally, Sec.~\ref{sec:conc} is devoted to conclusions.

\section{10D magnetized SYM in ${\cal N}=1$ superspace}
\label{sec:10d}

In this section, we briefly review a compactification of
10D SYM theory on flat 4D Minkowski spacetime times a
product of factorizable three tori $T^2 \times T^2 \times T^2$.
Then we derive a superfield description suitable for
such a compactification with certain magnetic fluxes
preserving 4D ${\cal N}=1$ supersymmetry.
The geometric (torus) parameter dependence is
explicitly shown in this procedure, which is important
to determine couplings between YM and moduli superfields
in the 4D effective action for zero-modes derived
in the next sections.

\subsection{Toroidal compactification of 10D SYM theory}
\label{ssec:10dsym}

The action of 10D SYM theory is given by
\begin{eqnarray}
S &=& \int d^{10}X\,\sqrt{-G}\,
\frac{1}{g^2}\,{\rm Tr}\left[
-\frac{1}{4}F^{MN}F_{MN}
+\frac{i}{2} \bar\lambda \Gamma^M D_M \lambda
\right],
\label{eq:10dsym}
\end{eqnarray}
where $g$ is a 10D YM gauge coupling constant
and the trace runs over the adjoint representation
of a gauge group. The 10D spacetime coordinates are
expressed by $X^M$, and the vector/tensor indices
$M,N=0,1,\ldots,9$ are lowered and raised by the 10D
metric $G_{MN}$ and its inverse $G^{MN}$, respectively.
The YM field strength $F_{MN}$ and
the covariant derivative $D_M$ are defined by
\begin{eqnarray}
F_{MN} &=&
\partial_M A_N - \partial_N A_M -i [A_M,A_N],
\nonumber \\
D_M \lambda &=& \partial_M \lambda -i [A_M,\lambda],
\nonumber
\end{eqnarray}
for a 10D vector (gauge) field $A_M$ and
a 10D Majorana-Weyl spinor field $\lambda$
satisfying
$\lambda^C = \lambda$ and
$\Gamma \lambda = +\lambda$
where
$\lambda^C$ is a 10D charge conjugation of $\lambda$,
and $\Gamma$ is a 10D chirality operator.

We decompose the 10D real coordinates $X^M=(x^\mu,y^m)$ into
4D Minkowski spacetime coordinates $x^\mu$ with $\mu=0,1,2,3$
and six dimensional (6D) extra space coordinates $y^m$ with
$m=4,\ldots,9$.
Note that $\mu=0$ describes the time component. Similarly
the 10D vector field is decomposed as $A_M=(A_\mu,A_m)$.
Our convention of the background metric is chosen as
\begin{eqnarray}
ds^2 &=& G_{NN}dX^MdX^N \ = \
\eta_{\mu \nu} dx^\mu dx^\nu + g_{mn} dy^m dy^n,
\nonumber
\end{eqnarray}
where
$\eta_{\mu \nu}={\rm diag}(-1,+1,+1,+1)$.
Then, we consider a torus compactification of internal
6D space $y^m$ by identifying $y^m \sim y^m + 2$.
The 6D torus is further assumed as a product of
factorizable three tori, $T^2 \times T^2 \times T^2$,
and then the extra 6D metric can be described as
\begin{eqnarray}
g_{mn} &=&
\begin{pmatrix}
g^{(1)} & 0 & 0 \\
0 & g^{(2)}& 0 \\
0 & 0 & g^{(3)}
\end{pmatrix},
\nonumber
\end{eqnarray}
where each of entries is a $2 \times 2$ matrix and
submatrices in the diagonal part are given by
\begin{eqnarray}
g^{(i)} &=& (2 \pi R_i)^2
\begin{pmatrix}
1 & {\rm Re}\,\tau_i \\
{\rm Re}\,\tau_i & |\tau_i|^2
\end{pmatrix},
\label{eq:gtorus}
\end{eqnarray}
for $i=1,2,3$.
The real and complex parameters $R_i$ and $\tau_i$,
respectively, determine the size and the shape of $i$th
torus $T^2$. Especially the area ${\cal A}^{(i)}$ of the
$i$th torus is described as
\begin{eqnarray}
{\cal A}^{(i)} &=& (2 \pi R_i)^2\,{\rm Im}\, \tau_i.
\nonumber
\end{eqnarray}

For convenience in a supersymmetric world, we introduce
complex (extra-dimensional) coordinates $z^i$ for $i=1,2,3$
and also complex vector components $A_i$ defined by
\begin{eqnarray}
z^i &\equiv& \frac{1}{2}(y^{2+2i} + \tau_i\,y^{3+2i}),
\qquad
\bar{z}^{\bar{i}} \ \equiv \ (z^i)^\ast,
\label{eq:cczi} \\
A_i &\equiv&
-\frac{1}{{\rm Im}\,\tau_i}
(\tau_i^\ast \,A_{2+2i}-A_{3+2i}),
\qquad
\bar{A}_{\bar{i}} \ \equiv \ (A_i)^\dagger.
\nonumber
\end{eqnarray}
The torus boundary conditions are given by
$z^i \sim z^i +1$ and $z^i \sim  z^i +\tau^i$.
A metric $h_{i\bar{j}}$ for the complex coordinates
is extracted from
\begin{eqnarray}
ds_{6D}^2 &=& g_{mn}dy^mdy^n
\ \equiv \ 2h_{i\bar{j}}dz^id\bar{z}^{\bar{j}},
\nonumber
\end{eqnarray}
and then we find
\begin{eqnarray}
h_{i\bar{j}} &=&
2\,(2 \pi R_i)^2\,\delta_{i\bar{j}}
\ = \ \delta_{{\rm i}\bar{\rm j}}\,
e_i^{\ {\rm i}}\,\bar{e}_{\bar{j}}^{\ \bar{\rm j}},
\nonumber
\end{eqnarray}
where $e_i^{\ {\rm i}}$ is a vielbein defined by
\begin{eqnarray}
e_i^{\ {\rm i}} &=&
\sqrt{2}\,(2 \pi R_i)\,\delta_i^{\ {\rm i}},
\nonumber
\end{eqnarray}
with its inverse $e_{{\rm i}}^{\ i}$
and complex conjugate
$\bar{e}_{\bar{i}}^{\ \bar{\rm i}}$
satisfying
\begin{eqnarray}
e_{{\rm i}}^{\ i} e_i^{\ {\rm j}}
&=& \delta_{{\rm i}}^{\ {\rm j}}, \qquad
e_i^{\ {\rm i}} e_{{\rm i}}^{\ j}
\ = \ \delta_i^{\ j}, \qquad
\bar{e}_{\bar{i}}^{\ \bar{\rm i}}
\ = \ (e_i^{\ {\rm i}})^\ast,
\nonumber
\end{eqnarray}
and Roman indices representing local Lorentz space.
The Italic (Roman) indices $i,j,\ldots$
(${\rm i}$, ${\rm j}, \ldots$) are lowered and raised
by the metric $h_{i\bar{j}}$
and its inverse $h^{\bar{i}j}$
($\delta_{{\rm i}\bar{\rm j}}$
and its inverse $\delta^{\bar{\rm i}{\rm j}}$),
respectively.

\subsection{Superfield description on nontrivial gauge background}
\label{ssec:sfss}

The 10D SYM theory has ${\cal N}=4$ supersymmetry in terms
of a 4D supercharge. The YM fields $A_M$ and $\lambda$ can
be decomposed into 4D ${\cal N}=1$ (on-shell) supermultiplets as
\begin{eqnarray}
{\bm V} &=& \left\{ A_\mu, \lambda_0 \right\}, \qquad
{\bm \phi}_i \ = \ \left\{ A_i, \lambda_i \right\}.
\nonumber
\end{eqnarray}
Here the 10D Majorana-Weyl spinor $\lambda$ is decomposed
into four 4D Weyl (or equivalently Majorana) spinors
$\lambda_0$ and $\lambda_i$ satisfying
\begin{eqnarray}
\Gamma^{(i)} \lambda_0 &=& +\lambda_0, \qquad
\Gamma^{(i)} \lambda_j \ = \ +\lambda_j \quad (i=j), \qquad
\Gamma^{(i)} \lambda_j \ = \ -\lambda_j \quad (i \ne j),
\label{eq:gamma5i}
\end{eqnarray}
where $\Gamma^{(i)}$ for each $i=1,2,3$ are a chirality
operator associated with 6D spacetime coordinates $(x^\mu,z^i)$.
Note that an eigenvalue of the 10D chirality operator $\Gamma$
is obtained by a product of the eigenvalues of
$\Gamma^{(1)}$, $\Gamma^{(2)}$ and $\Gamma^{(3)}$.
If we write the chirality of $i$th complex coordinate $z_i$
in the $i$th subscript of $\lambda$ like
$\lambda_{\pm \pm \pm}$,
the decomposed spinor fields $\lambda_0$ and $\lambda_i$
are identified with the chirality eigenstate
$\lambda_{\pm \pm \pm}$ as
\begin{eqnarray}
\lambda_0 &=& \lambda_{+++}, \qquad
\lambda_1 \ = \ \lambda_{+--}, \qquad
\lambda_2 \ = \ \lambda_{-+-}, \qquad
\lambda_3 \ = \ \lambda_{--+}.
\nonumber
\end{eqnarray}
Note that the components
$\lambda_{---}$, $\lambda_{-++}$,
$\lambda_{+-+}$ and $\lambda_{++-}$
do not exist in the 10D Majorana-Weyl spinor $\lambda$
due to the condition $\Gamma \lambda = +\lambda$.

The above ${\cal N}=1$ vector multiplet ${\bm V}$ and
chiral multiplets ${\bm \phi}_i$ are expressed, respectively,
by a vector superfield $V$ and a chiral superfield $\phi_i$.
Our definition of superfields is as follows:
\begin{eqnarray}
V &\equiv& -\theta\sigma^\mu\bar\theta A_\mu
+i\bar\theta\bar\theta\theta\lambda_0
-i\theta\theta\bar\theta\bar\lambda_0
+\frac12\theta\theta\bar\theta\bar\theta D,
\nonumber \\
\phi_i &\equiv& \frac1{\sqrt2} A_i
+\sqrt2\theta\lambda_i+\theta\theta F_i,
\nonumber
\end{eqnarray}
where $\theta$ and $\bar\theta$ are
Grassmann coordinates of 4D ${\cal N}=1$ superspace with
spinor indices $\alpha$ and $\dot\alpha$ omitted
respectively\footnote{Our conventions mostly
follow those in Ref.~\cite{Wess:1992cp}.}.
The superfield description allows us to express the
10D SYM action (\ref{eq:10dsym}) in ${\cal N}=1$
superspace as~\cite{ArkaniHamed:2001tb}
\begin{eqnarray}
S &=& \int d^{10}X \sqrt{-G} \left[ \int d^4\theta\,{\cal K}
+\left\{ \int d^2 \theta\,\left(
\frac{1}{4g^2} {\cal W}^\alpha {\cal W}_\alpha
+ {\cal W} \right) +{\rm h.c.} \right\}
\right],
\label{eq:10dss}
\end{eqnarray}
where
\begin{eqnarray}
{\cal K} &=& \frac{2}{g^2}  h^{\bar{i}j}
{\rm Tr} \left[\left(\sqrt2\bar\partial_{\bar{i}}
+\bar\phi_{\bar{i}}\right)e^{-V}
\left(-\sqrt2\partial_j + \phi_j \right)e^V
+\bar\partial_{\bar{i}} e^{-V} \partial_j e^V \right]
+{\cal K}_{\rm WZW},
\nonumber \\
{\cal W} &=& \frac{1}{g^2}
\epsilon^{{\rm i}{\rm j}{\rm k}}
e_{{\rm i}}^{\ i} e_{{\rm j}}^{\ j} e_{{\rm k}}^{\ k}
{\rm Tr} \left[ \sqrt{2}\, \phi_i
\left(\partial_j\phi_k-\frac{1}{3\sqrt2}
\left[\phi_j,\phi_k\right]\right)\right],
\nonumber
\end{eqnarray}
with a totally antisymmetric tensor
$\epsilon^{{\rm i}{\rm j}{\rm k}}$
and $\epsilon^{123}=1$.
The field strength superfield ${\cal W}_\alpha$ is defined by
${\cal W}_\alpha \equiv
-\frac{1}{4} \bar{D} \bar{D} e^{-V} D_\alpha e^V$
where $D_\alpha$ and $\bar{D}_{\dot\alpha}$
are a supercovariant derivative and its conjugate with
4D spinor indices $\alpha$ and $\dot\alpha$, respectively.
The term ${\cal K}_{\rm WZW}$ corresponds to a
Wess-Zumino-Witten term
which vanishes in Wess-Zumino (WZ) gauge.
The equations of motion for
auxiliary fields $D$ and $F_i$ lead to
\begin{eqnarray}
D &=& - h^{\bar{i}j} \left(
\bar\partial_{\bar{i}} A_j + \partial_j \bar{A}_{\bar{i}}
+\frac{1}{2} \left[ \bar{A}_{\bar{i}}, A_j \right] \right),
\label{eq:osd} \\
\bar{F}_{\bar{i}} &=& -h_{j\bar{i}}\,
\epsilon^{{\rm j}{\rm k}{\rm l}}
e_{{\rm j}}^{\ j} e_{{\rm k}}^{\ k} e_{{\rm l}}^{\ l}
\left( \partial_k A_l
-\frac{1}{4} \left[ A_k,\, A_l \right] \right).
\label{eq:osfi}
\end{eqnarray}
The original SYM action (\ref{eq:10dsym}) is obtained
after integrating over the superspace in Eq.~(\ref{eq:10dss})
and substituting on-shell values
(\ref{eq:osd}) and (\ref{eq:osfi}).

We assume 4D Lorentz invariant and (at least ${\cal N}=1$)
supersymmetric VEVs of fields,
\begin{eqnarray}
\vev{A_i} &\ne& 0, \qquad
\vev{A_\mu} \ = \
\vev{\lambda_0} \ = \
\vev{\lambda_i} \ = \
\vev{F_i} \ = \
\vev{D} \ = \ 0.
\label{eq:vevs}
\end{eqnarray}
We will see later\footnote{See
Eqs.(\ref{eq:abd}), (\ref{eq:abfi}) and (\ref{eq:susy}).}
that certain magnetized background satisfy Eq.~(\ref{eq:vevs}).
Then we extract fluctuations $\tilde{V}$ and $\tilde{\phi}_i$
of superfields $V$ and $\phi_i$, respectively, around the
vacuum configuration (\ref{eq:vevs}) as
\begin{eqnarray}
V &\equiv& \vev{V} + \tilde V, \qquad
\phi_i \ \equiv \ \vev{\phi_i} + \tilde\phi_i,
\nonumber
\end{eqnarray}
where $\vev{V}=0$ and $\vev{\phi_i}=\vev{A_i}/\sqrt{2}$
due to Eq.~(\ref{eq:vevs}).
For a notational convenience, in the following,
we omit tildes and use original notation $V$ and $\phi_i$
for their corresponding fluctuations around the vacuum.
Then, in the WZ-gauge, the functions
${\cal K}$ and ${\cal W}$ are expanded in powers of $V$ as
\begin{eqnarray}
{\cal K} &=& \frac{2}{g^2} h^{\bar{i}j} {\rm Tr} \Bigg[
\bar\phi_{\bar{i}} \phi_j
+\sqrt{2} \left\{ \left(\bar\partial_{\bar{i}} \phi_j
+\frac{1}{\sqrt2}
\left[ \vev{\bar\phi_{\bar{i}}},\,\phi_j \right]
+{\rm h.c.} \right)
+\frac{1}{\sqrt2}
\left[ \bar\phi_{\bar{i}},\,\phi_j \right] \right\} V
\nonumber \\ && \qquad \qquad
+(\bar\partial_{\bar{i}}V)(\partial_j V)
+ \frac{1}{2} \left( \bar\phi_{\bar{i}} \phi_j
+\phi_j \bar\phi_{\bar{i}} \right) V^2
-\bar\phi_{\bar{i}} V \phi_j V \Bigg]
+{\cal K}^{({\rm D})}
+{\cal K}^{({\rm br})},
\label{eq:kfv} \\
{\cal W} &=& \frac{1}{g^2} \epsilon^{{\rm i}{\rm j}{\rm k}}
e_{{\rm i}}^{\ i} e_{{\rm j}}^{\ j} e_{{\rm k}}^{\ k}
{\rm Tr} \left[ \sqrt{2} \left(
\partial_i\phi_j -\frac{1}{\sqrt{2}}
\left[ \vev{\phi_i},\,\phi_j \right] \right) \phi_k
-\frac{2}{3} \phi_i \phi_j \phi_k \right]+{\cal W}^{(\rm{F})},
\label{eq:wfv}
\end{eqnarray}
where
\begin{eqnarray}
{\cal K}^{({\rm br})} &=&
\frac{1}{g^2} h^{\bar{i}j} {\rm Tr} \Bigg[
\frac{1}{2} \left(
\vev{\bar\phi_{\bar{i}}} \vev{\phi_j}
+\vev{\phi_j} \vev{\bar\phi_{\bar{i}}}
\vev{\bar\phi_{\bar{i}}} \phi_j
+\phi_j \vev{\bar\phi_{\bar{i}}}
+\bar\phi_{\bar{i}} \vev{\phi_j}
+\vev{\phi_j} \bar\phi_{\bar{i}}
\right) V^2
\nonumber \\ && \qquad \qquad
-\vev{\bar\phi_{\bar{i}}} V \vev{\phi_j} V
-\vev{\bar\phi_{\bar{i}}} V \phi_j V
-\bar\phi_{\bar{i}} V \vev{\phi_j} V
 \Bigg],
\label{eq:unb}
\end{eqnarray}
and we omit terms in higher-powers of
$\theta$ and $\bar\theta$ than
$\theta^2 \bar\theta^2$ which vanish
in the superspace action (\ref{eq:10dss}).
The two terms of ${\cal K}^{({\rm D})}$ and ${\cal W}^{(\rm{F})}$
are eliminated in the case that
the SUSY conditions (\ref{eq:osd}) and (\ref{eq:osfi}) are satisfied.
The vacuum configuration (\ref{eq:vevs}) in general
breaks the gauge symmetry (as well as the
higher-dimensional supersymmetry) of original SYM theory
down to its subgroup, and vector multiplets associated
with the broken symmetries become massive. Such masses
are generated in ${\cal K}^{({\rm br})}$.

\subsection{Zero-mode equations in superspace}
\label{ssec:zme}

On the ${\cal N}=1$ supersymmetric background (\ref{eq:vevs}),
Kaluza-Klein (KK) mode-equations should be written as superfield
equations. Then, mode expansions of superfields $V$ and $\phi_j$
are expressed as
\begin{eqnarray}
V(x^\mu,{\bm z},\bar{\bm z})
&=& \sum_{{\bm n}} \left( \prod_i
f_0^{(i),n_i}(z^i,\bar{z}^{\bar{i}}) \right)
V^{{\bm n}}(x^\mu),
\label{eq:mev} \\
\phi_j(x^\mu,{\bm z},\bar{\bm z})
&=& \sum_{{\bm n}} \left( \prod_i
f_j^{(i),n_i}(z^i,\bar{z}^{\bar{i}}) \right)
\phi_j^{{\bm n}}(x^\mu),
\label{eq:mephi}
\end{eqnarray}
where
${\bm z}=(z^1,z^2,z^3)$,
$\bar{\bm z}=(
\bar{z}^{\bar{1}},\bar{z}^{\bar{2}},\bar{z}^{\bar{3}})$,
${\bm n} = (n_1, n_2, n_3)$
and $n_i \in {\bf Z}$ for $i,j=1,2,3$.
The 4D fields $V^{{\bm n}}$ and $\phi_j^{{\bm n}}$
are KK modes of 10D fields $V$ and $\phi_j$, respectively,
with KK momenta of three tori labeled by ${\bm n}$.
The functions $f_0^{(i),n_i}$ and $f_j^{(i),n_i}$ are
wavefunctions of $V^{{\bm n}}$ and $\phi_j^{{\bm n}}$,
respectively, in the $i$th torus.
Note that all the quantities
$\{V, \phi_j, V^{{\bm n}}, \phi_j^{{\bm n}},
f_0^{(i),n_i}, f_j^{(i),n_i} \}={\cal G}$
in Eqs.~(\ref{eq:mev}) and (\ref{eq:mephi})
have the same subscripts of YM indices implicitly,
e.g.,
${\cal G}={\cal G}_{AB}$ with $A,B=1,2,\ldots,N$
for $U(N)$ SYM theory,
because the both-hand sides of
Eqs.~(\ref{eq:mev}) and (\ref{eq:mephi})
are adjoint matrices of YM gauge group,
and wavefunctions can be different from each other
element by element due to a possible gauge symmetry
breaking caused by $\vev{A_i} \ne 0$ in Eq.~(\ref{eq:vevs}).
We should remark that $V^{{\bm n}}$ and $\phi_j^{{\bm n}}$
are a vector and chiral superfields, respectively, while
$f_0^{(i),n_i}$ and $f_j^{(i),n_i}$ are independent of
superspace coordinates $\theta$ and $\bar\theta$ on a
${\cal N}=1$ supersymmetric background (\ref{eq:vevs}).

Substituting Eqs.~(\ref{eq:mev}) and (\ref{eq:mephi})
into the superspace action (\ref{eq:10dss})
with Eqs.~(\ref{eq:kfv}) and (\ref{eq:wfv}),
we find zero-mode equations as
\begin{eqnarray}
\partial_i f_0^{(i),n_i=0} -\frac1{\sqrt2}
\left[ \vev{\phi_i},\,f_0^{(i),n_i=0} \right]
&=& 0,
\nonumber \\
\bar\partial_{\bar{i}} f_j^{(i),n_i=0} +\frac1{\sqrt2}
\left[ \vev{\bar\phi_{\bar{i}}},\,f_j^{(i),n_i=0} \right]
&=& 0 \qquad (i=j),
\label{eq:kkeii} \\
\partial_i f_j^{(i),n_i=0} -\frac1{\sqrt2}
\left[ \vev{\phi_i},\,f_j^{(i),n_i=0} \right]
&=& 0 \qquad (i \ne j),
\label{eq:kkeij}
\end{eqnarray}
with which terms quadratic in superfields $\phi_i$ and $V$
are eliminated in Eqs.~(\ref{eq:kfv}) and (\ref{eq:wfv})
except in Eq.~(\ref{eq:unb}), and then vanishing masses
for zero-modes $\phi_j^{{\bm n}={\bm 0}}$ are guaranteed.
Note that a sign-difference between the second terms in the
left-hand sides of Eqs.~(\ref{eq:kkeii}) and (\ref{eq:kkeij})
comes from the chirality structure (\ref{eq:gamma5i}).
A remarkable point is that zero-mode equations are written
for superfields. Such a superfield description allows us
to perform a dimensional reduction while keeping a
manifest ${\cal N}=1$ superspace structure preserved by
the gauge background (\ref{eq:vevs}).

\section{4D effective action}
\label{sec:4d}

In the following, we focus on 4D massless modes (called zero-modes)
in chiral superfields $\phi_i$ which possess nontrivial wavefunction
profiles under the existence of magnetic fluxes in extra dimensions.
We derive a 4D effective action for zero-modes by solving the zero-mode
equations and performing integrations over extra space coordinates
in the 10D action, while keeping the ${\cal N}=1$ superspace structure
preserved by the flux background.

\subsection{Dimensional reduction with magnetic fluxes}
\label{ssec:dimred}

Because we focus on zero-modes in the following,
for a notational convenience, their wavefunctions are denoted as
\begin{eqnarray}
f_0^{(i),n_i=0} &\equiv& f_0^{(i)}, \qquad
f_j^{(i),n_i=0} \ \equiv \ f_j^{(i)}.
\nonumber
\end{eqnarray}
After substituting $\vev{\phi_i}=\vev{A_i}/\sqrt{2}$
into Eqs.~(\ref{eq:kkeii}) and (\ref{eq:kkeij}),
zero-mode equations are expressed by
\begin{eqnarray}
\bar\partial_{\bar{i}} f_j^{(i)} +\frac{1}{2}
\left[ \vev{\bar{A}_{\bar{i}}},\,f_j^{(i)} \right]
&=& 0 \qquad (i=j),
\label{eq:zmeii} \\
\partial_i f_j^{(i)} -\frac{1}{2}
\left[ \vev{A_i},\,f_j^{(i)} \right]
&=& 0 \qquad (i \ne j).
\label{eq:zmeij}
\end{eqnarray}
It is found that Eqs.(\ref{eq:zmeii}) and (\ref{eq:zmeij})
are equivalent to zero-mode equations for charged fermion
fields $\lambda_j$ in the gauge background~\cite{Cremades:2004wa}
as it should be.

In the following we consider the case that the YM gauge group is
$U(N)$ and assume a supersymmetric (Abelian) magnetic background:
\begin{eqnarray}
\vev{A_i} &=& \frac{\pi}{{\rm Im}\, \tau_i}
\left( M^{(i)}\,\bar{z}_{\bar{i}}+\bar\zeta_i \right),
\nonumber
\end{eqnarray}
where
\begin{eqnarray}
M^{(i)} &=& {\rm diag}
(M_1^{(i)},M_2^{(i)},\ldots,M_N^{(i)}),
\nonumber \\
\zeta_i &=& {\rm diag}
(\zeta_1^{(i)},\zeta_2^{(i)},\ldots,\zeta_N^{(i)}),
\nonumber
\end{eqnarray}
are $N \times N$ diagonal matrices of Abelian
magnetic fluxes and Wilson-lines, respectively.
Here, the magnetic fluxes must satisfy
$M_1^{(i)},M_2^{(i)},\ldots,M_N^{(i)} \in {\bf Z}$
due to the Dirac's quantization condition.
We assume that the above magnetic fluxes
are further constrained in such a way
that they satisfy supersymmetry condition
\begin{eqnarray}
h^{\bar{i}j} \left( \bar\partial_{\bar{i}} \vev{A_j}
+ \partial_j \vev{\bar{A}_{\bar{i}}} \right) &=& 0,
\label{eq:abd} \\
\epsilon^{{\rm j}{\rm k}{\rm l}}
e_{{\rm k}}^{\ k} e_{{\rm l}}^{\ l}
\partial_k \vev{A_l} &=& 0,
\label{eq:abfi}
\end{eqnarray}
in order to satisfy Eq.~(\ref{eq:vevs}) with
Eqs.~(\ref{eq:osd}) and (\ref{eq:osfi})
for the Abelian background.

In the case that all the magnetic fluxes
$M_1^{(i)},M_2^{(i)},\ldots,M_N^{(i)}$ take different values
from each other, the YM gauge symmetry is broken down as
$U(N) \to U(1)^N$ by the existence of fluxes.
On the other hand, if any of them degenerate, e.g.,
\begin{eqnarray}
&M_1^{(i)}=M_2^{(i)}=\cdots=M_{N_1}^{(i)},&
\nonumber \\
&M_{N_1+1}^{(i)}=M_{N_1+2}^{(i)}=\cdots=M_{N_1+N_2}^{(i)},&
\nonumber \\
&\vdots&
\nonumber \\
&M_{N_1+N_2+\cdots+N_{\tilde{N}-1}+1}^{(i)}
=M_{N_1+N_2+\cdots+N_{\tilde{N}-1}+2}^{(i)}=\cdots
=M_{N_1+N_2+\cdots+N_{\tilde{N}-1}+N_{\tilde{N}}}^{(i)},&
\label{eq:gsb}
\end{eqnarray}
with $\sum_a N_a=N$ and $M_{N_a} \ne M_{N_b}$
for $a,b=1,2,\ldots,\tilde{N}$ and $a \ne b$,
the breaking pattern is changed as
$U(N) \to \prod_a U(N_a)$.
The same holds for Wilson-lines.
In the following, indices $a,b,c=1,2,\ldots,\tilde{N}$ label the
unbroken YM subgroups on the flux and Wilson-line background,
and traces in expressions are performed within such subgroups.

{}From Eqs.
(\ref{eq:zmeii}) and (\ref{eq:zmeij}),
for the unbroken YM subgroup labeled by $a$ and $b$,
we find zero-mode equations
$(f_j^{(i)})_{ab}$ as
\begin{eqnarray}
\left[ \bar\partial_{\bar{i}} +\frac{\pi}{2{\rm Im}\,\tau_i}
\left( M^{(i)}_{ab}\,z_i
+\zeta^{(i)}_{ab} \right) \right] (f_j^{(i)})_{ab}
&=& 0 \qquad (i=j),
\label{eq:czmeii} \\
\left[ \partial_i -\frac{\pi}{2{\rm Im}\,\tau_i}
\left( M^{(i)}_{ab}\,\bar{z}_{\bar{i}}
+\bar\zeta^{(i)}_{ab} \right) \right] (f_j^{(i)})_{ab}
&=& 0 \qquad (i \ne j),
\label{eq:czmeij}
\end{eqnarray}
where
\begin{eqnarray}
M^{(i)}_{ab} &=& M^{(i)}_{N_a}-M^{(i)}_{N_b}, \qquad
\zeta^{(i)}_{ab}=\zeta^{(i)}_{N_a}-\zeta^{(i)}_{N_b}.
\nonumber
\end{eqnarray}

For $i=j$, a normalizable solution of
Eq.~(\ref{eq:czmeii}) is found as
\begin{eqnarray}
(f_j^{(i)})_{ab} &=&
f^{I^{(i)}_{ab}} \ \equiv \
\left\{ \begin{array}{ll}
\displaystyle
\Theta^{I^{(i)}_{ab},\,M^{(i)}_{ab}}
\left( \tilde{z}_i \right)
& \quad (M^{(i)}_{ab} > 0) \\
\displaystyle ({\cal A}^{(i)})^{-1/2} &
\quad (M^{(i)}_{ab} = 0) \\
\displaystyle
0 & \quad (M^{(i)}_{ab} < 0)
\end{array} \right.,
\label{eq:fijab}
\end{eqnarray}
where
$\tilde{z}_i \equiv
z_i + \frac{\zeta^{(i)}_{ab}}{M^{(i)}_{ab}}$ and
\begin{eqnarray}
I^{(i)}_{ab} &=&
\left\{ \begin{array}{ll}
1,\, \ldots,\, \left| M^{(i)}_{ab} \right|
& \quad (M^{(i)}_{ab} > 0) \\
0 & \quad (M^{(i)}_{ab} = 0) \\
{\rm no \ solution} & \quad (M^{(i)}_{ab} < 0)
\end{array} \right..
\label{eq:igen}
\end{eqnarray}
Due to the effect of chirality projection by fluxes,
no zero-mode appears for $M^{(i)}_{ab} < 0$.
On the other hand, for  $M^{(i)}_{ab} > 0$, there
appear $M^{(i)}_{ab}$ zero-modes, and these zero-modes are
labeled by the index $I^{(i)}_{ab}$.
In Eq.~(\ref{eq:fijab}), the wavefunction profile
$\Theta^{I,\,M}(z)$ is determined as
\begin{eqnarray}
\Theta^{I,\,M}(z) &\equiv&
{\cal N}_M \, e^{\pi i
\frac{{\rm Im}\,z}{{\rm Im}\,\tau} M z}\,
\vartheta
\begin{bmatrix} \frac{I}{M} \\ 0 \end{bmatrix}
\left( M z, M \tau \right),
\nonumber
\end{eqnarray}
where $\vartheta$ represents the
Jacobi theta-function:
\begin{eqnarray}
\vartheta
\begin{bmatrix} a \\ b \end{bmatrix}
\left( \nu, \tau \right) &=&
\sum_{l \in {\bf Z}}
e^{\pi i \left( a + l \right)^2 \tau}
e^{2 \pi i \left( a + l \right)
\left( \nu + b \right)} .
\nonumber
\end{eqnarray}
The normalization constant ${\cal N}_M$ is found as
\begin{eqnarray}
{\cal N}_{M^{(i)}} &=& \left(
\frac{2\,{\rm Im}\,\tau_i\,|M^{(i)}|}{({\cal A}^{(i)})^2}
\right)^{1/4},
\nonumber
\end{eqnarray}
from a normalization condition,
\begin{eqnarray}
\int dz_i d \bar{z}_i \sqrt{{\rm det}\,g^{(i)}}\,
f^I\,(f^J)^\ast &=& \delta^{IJ},
\nonumber
\end{eqnarray}
for $I,J \ne 0$ or $I=J=0$.

For $i \ne j$, on the other hand,
from Eq.~(\ref{eq:czmeij}) we find
\begin{eqnarray}
(f_j^{(i)})_{ab} &=&
f^{I^{(i)}_{ab}} \ \equiv \
\left\{ \begin{array}{ll}
\displaystyle
0 & \quad (M^{(i)}_{ab} > 0) \\
\displaystyle ({\cal A}^{(i)})^{-1/2} &
\quad (M^{(i)}_{ab} = 0) \\
\displaystyle
\left( \Theta^{I^{(i)}_{ab},\,|M^{(i)}_{ab}|}
\left( \tilde{z}_i \right) \right)^\ast
& \quad (M^{(i)}_{ab} < 0)
\end{array} \right.,
\nonumber
\end{eqnarray}
where
$\tilde{z}_i \equiv
z_i - \frac{\zeta^{(i)}_{ab}}{|M^{(i)}_{ab}|}$,
instead of Eq.~(\ref{eq:fijab}).

\subsection{4D effective action for zero-modes}
\label{ssec:zm4deff}

Due to the gauge symmetry breaking
$U(N) \to \prod_a U(N_a)$ with $a=1,\ldots,\tilde{N}$
caused by the fluxes satisfying Eq.~(\ref{eq:gsb}),
off-diagonal elements $(V^{{\bm n}={\bm 0}})_{ab}$
($a \ne b$) obtain mass terms in Eq.~(\ref{eq:unb}),
while diagonal elements
$(V^{{\bm n}={\bm 0}})_{aa}$ do not.
Then, we express the zero-modes
$(V^{{\bm n}={\bm 0}})_{aa}$,
which contain gauge fields for the unbroken gauge symmetry
$\prod_a U(N_a)$, as
\begin{eqnarray}
(V^{{\bm n}={\bm 0}})_{aa} &\equiv& V^a.
\nonumber
\end{eqnarray}
On the other hand,
from Eq.~(\ref{eq:igen}),
for $^\exists j \ne i$ with
$M^{(j)}_{ab} < 0$ and $M^{(i)}_{ab} > 0$
we find the zero-mode field $(\phi_j^{{\bm n}={\bm 0}})_{ab}$
degenerates with a total degeneracy
$N_{ab}=\prod_i \left| M^{(i)}_{ab} \right|$,
while $(\phi_j^{{\bm n}={\bm 0}})_{ba}$ has no zero-mode solution,
yielding 4D chiral generations in the $ab$-sector.
The opposite is true for
$M^{(j)}_{ab} > 0$ and $M^{(i)}_{ab} < 0$
yielding 4D chiral generations in the $ba$-sector.
Therefore, we denote the zero-modes
$(\phi_j^{{\bm n}={\bm 0}})_{ab}$
with the degeneracy $N_{ab}$ as
\begin{eqnarray}
(\phi_j^{{\bm n}={\bm 0}})_{ab} &\equiv&
g\,\phi_j^{{\cal I}_{ab}},
\nonumber
\end{eqnarray}
where ${\cal I}_{ab} \equiv
(I^{(1)}_{ab}, I^{(2)}_{ab}, I^{(3)}_{ab})$ labels
the degeneracy (generations).
We normalize $\phi_j^{{\cal I}_{ab}}$ by
the 10D YM coupling constant $g$.

The analytic expressions of zero-mode wavefunctions
allow us to derive 4D effective action after substituting
the mode expansion (\ref{eq:mephi})
and integrating 10D action (\ref{eq:10dss})
over 6D extra space coordinates $z_i$ and $\bar{z}^{\bar{i}}$.
The effective action for zero-mode chiral superfields $\phi_j^{{\cal
I}_{ab}}$ is then found as
\begin{eqnarray}
S_{\rm eff} &=&
\int d^4x \left[ \int d^4 \theta\,{\cal K}_{\rm eff}
+\left\{ \int d^2 \theta \left(
\sum_a \frac{1}{4g_a^2} \,W^{a,\alpha} W^a_\alpha
+{\cal W}_{\rm eff} \right)
+{\rm h.c.} \right\} \right],
\label{eq:seff}
\end{eqnarray}
where
\begin{eqnarray}
{\cal K}_{\rm eff} &=& \sum_{i,j}\,\sum_{a,b}\,
\sum_{{\cal I}_{ab}}
\tilde{Z}_{{\cal I}_{ab}}^{\bar{i}j} {\rm tr} \left[
\bar\phi_{\bar{i}}^{{\cal I}_{ab}} e^{-V^a}
\phi_j^{{\cal I}_{ab}} e^{V^b} \right],
\nonumber \\
{\cal W}_{\rm eff} &=&
\sum_{i,j,k}\,\sum_{a,b,c}\,
\sum_{{\cal I}_{ab},\,
{\cal I}_{bc},\,
{\cal I}_{ca}}
\tilde\lambda_{{\cal I}_{ab}
{\cal I}_{bc}
{\cal I}_{ca}}^{ijk}
{\rm tr} \left[
\phi_i^{{\cal I}_{ab}}
\phi_j^{{\cal I}_{bc}}
\phi_k^{{\cal I}_{ca}} \right],
\nonumber
\end{eqnarray}
and
\begin{eqnarray}
W_\alpha^a &=& -\frac{1}{4} \bar{D} \bar{D}
e^{-V^a} D_\alpha e^{V^a}, \qquad
g_a \ = \ g \left( \prod_i {\cal A}^{(i)} \right)^{-1/2}.
\nonumber
\end{eqnarray}
The K\"ahler metric $\tilde{Z}_{{\cal I}_{ab}}$
and the holomorphic Yukawa couplings
$\tilde\lambda_{{\cal I}_{ab}
{\cal I}_{bc}
{\cal I}_{ca}}$
can be written as
\begin{eqnarray}
\tilde{Z}_{{\cal I}_{ab}}^{\bar{i}j}
&=& 2 h^{\bar{i}j},
\label{eq:km} \\
\tilde\lambda_{
{\cal I}_{ab}
{\cal I}_{bc}
{\cal I}_{ca}}^{ijk} &=&
-\frac{2g}{3}
\epsilon^{{\rm i}{\rm j}{\rm k}}
e_{{\rm i}}^{\ i}
e_{{\rm j}}^{\ j}
e_{{\rm k}}^{\ k}
\prod_r
\tilde\lambda^{(r)}_{
I^{(r)}_{ab}I^{(r)}_{bc}I^{(r)}_{ca}},
\label{eq:yc}
\end{eqnarray}
where $r=1,2,3$ and
\begin{eqnarray}
\tilde\lambda^{(r)}_{
I^{(r)}_{ab}I^{(r)}_{bc}I^{(r)}_{ca}}
&=& \int d^2z^r \sqrt{{\rm det}\,g^{(r)}}\,
f^{I^{(r)}_{ab}}f^{I^{(r)}_{bc}}f^{I^{(r)}_{ca}}.
\label{eq:intyci}
\end{eqnarray}

For $a$, $b$ and $c$ satisfying
$M^{(r)}_{ab}M^{(r)}_{bc}M^{(r)}_{ca}>0$
(that is equivalent to
$M^{(r)}_{ac}M^{(r)}_{cb}M^{(r)}_{ba}<0$),
the overlap integral (\ref{eq:intyci}) is evaluated as
\begin{eqnarray}
\tilde\lambda^{(r)}_{
I^{(r)}_{ab}I^{(r)}_{bc}I^{(r)}_{ca}} &=&
\left\{ \begin{array}{lll}
\tilde\lambda^{(r)}_{ab,c}
&& (M^{(r)}_{ab}>0) \\*[5pt]
\tilde\lambda^{(r)}_{bc,a}
&& (M^{(r)}_{bc}>0) \\*[5pt]
\tilde\lambda^{(r)}_{ca,b}
&& (M^{(r)}_{ca}>0)
\end{array} \right.,
\label{eq:yci}
\end{eqnarray}
where
\begin{eqnarray}
\tilde\lambda^{(r)}_{ab,c} &=&
{\cal N}_{M^{(r)}_{ab}}^{-1}
{\cal N}_{M^{(r)}_{bc}}
{\cal N}_{M^{(r)}_{ca}}
\sum_{m=1}^{M^{(r)}_{ab}}
\delta_{I^{(r)}_{bc}+I^{(r)}_{ca}
-m M^{(r)}_{bc},\, I^{(r)}_{ab}}\,
\nonumber \\ && \times
\exp \left[ \frac{\pi i}{{\rm Im}\, \tau_r}
\left(
\frac{\bar\zeta^{(r)}_{ab}}{M^{(r)}_{ab}}
{\rm Im}\,\zeta^{(r)}_{ab}
+\frac{\bar\zeta^{(r)}_{bc}}{M^{(r)}_{bc}}
{\rm Im}\,\zeta^{(r)}_{bc}
+\frac{\bar\zeta^{(r)}_{ca}}{M^{(r)}_{ca}}
{\rm Im}\,\zeta^{(r)}_{ca}
\right) \right]
\nonumber \\ && \times
\vartheta
\begin{bmatrix}
\frac{
M^{(r)}_{bc}I^{(r)}_{ca}
-M^{(r)}_{ca}I^{(r)}_{bc}
+m M^{(r)}_{bc}M^{(r)}_{ca}}{
M^{(r)}_{ab}M^{(r)}_{bc}M^{(r)}_{ca}}
\\ 0 \end{bmatrix} \left(
\bar\zeta^{(r)}_{ca}M^{(r)}_{bc}
-\bar\zeta^{(r)}_{bc}M^{(r)}_{ca},\,
-\bar\tau_r M^{(r)}_{ab}M^{(r)}_{bc}M^{(r)}_{ca}
\right).
\nonumber
\end{eqnarray}
For $a$, $b$ and $c$ satisfying
$M^{(r)}_{ab}M^{(r)}_{bc}M^{(r)}_{ca}=0$,
on the other hand,
the integral (\ref{eq:intyci}) is given by
$\tilde\lambda^{(r)}_{
I^{(r)}_{ab}I^{(r)}_{bc}I^{(r)}_{ca}}
= ({\cal A}^{(r)})^{-1/2}$
instead of Eq.~(\ref{eq:yci}).

\section{Local supersymmetry and moduli multiplets}
\label{sec:sugra}

The above derivation of 4D effective action has
been performed in a limit of global supersymmetry,
because our staring point is 10D SYM theory.
{}From both theoretical and phenomenological
viewpoints, however, theories with a local
supersymmetry are desirable. It is well known that
10D SYM theories can be embedded into supergravity.
Actually, low energy effective theories of
heterotic and type I superstrings as well as
type II orientifold/D-brane models are categorized
into such a supergravity-YM system.

The existence of local supersymmetry implies that
SYM system is coupled to gravity. With nonvanishing
gravitational interactions, theories in
higher-dimensional spacetime in general yield more
massless modes from gravitational fields which form
supermultiplets in the 4D effective theories.
Especially, massless modes originating from the
extra-dimensional components of higher-dimensional
tensor (such as 10D graviton fields) and vector fields
are called moduli which form chiral multiplets in 4D
${\cal N}=1$ supersymmetry.

In the following, we assume that there exists a local
supersymmetry at the starting point of our previous
analysis. Then, we show how to recover the local
supersymmetry in the 4D effective theory, and identify
the dependence of the effective action on geometric
(so-called closed string) moduli
as well as dilaton superfields.

\subsection{4D ${\cal N}=1$ effective supergravity}
\label{ssec:4dsugra}

The action for
4D ${\cal N}=1$ conformal supergravity~\cite{Kaku:1977rk}
is generally written as\footnote{
The superspace integrals
$\int d^4 \theta \cdots$ and  $\int d^2 \theta \cdots$
in Eq.~(\ref{eq:csugra}) are a kind of simplified expressions
for a notational convenience, which should be interpreted as
$D$-term formula $[\ldots]_D$ and
$F$-term formula $[\ldots]_F$ of
superconformal tensor calculus~\cite{Kugo:1982cu}, respectively.}
\begin{eqnarray}
S_{{\cal N}=1} &=& \int d^4x \sqrt{-g^C} \Bigg[
-3 \int d^4 \theta\,\bar{C}C \, e^{-K/3}
\nonumber \\ && \qquad \qquad \qquad
+\left\{ \int d^2 \theta \left(
\frac{1}{4} \sum_a f_a\, W^{a,\alpha} W^a_\alpha
+C^3 W\right) +{\rm h.c.} \right\} \Bigg],
\label{eq:csugra}
\end{eqnarray}
where $C$ is a compensator chiral superfield, whose
lowest component $C|_0 \equiv C|_{\theta=\bar\theta=0}$
in the $\theta$ and $\bar\theta$ expansion relates the
4D metric $g^C_{\mu \nu}$ in Eq.~(\ref{eq:csugra})
with the one in Einstein frame $g^E_{\mu \nu}$ as
$g^C_{\mu \nu} =
(C \bar{C}|_0)^{-1} e^{K|_0/3} g^E_{\mu \nu}$.
Here and hereafter we denote the lowest component
of a (function of) superfield $\Phi$ in the $\theta$ and
$\bar\theta$ expansion as
$\Phi|_{\theta=\bar\theta=0} \equiv \Phi|_0$.
The action of Poincare supergravity in Einstein frame
is obtained~\cite{Kaku:1978ea}
by a dilatation gauge fixing $C|_0 = e^{K|_0/6}$
in the conformal supergravity action~(\ref{eq:csugra}).
Here and hereafter, we work in a unit that
the 4D Planck scale is unity.

When we consider a situation that our starting 10D SYM theory
is embedded in 10D supergravity, it is important to notify
that there exists a scalar field $\phi_{10}$ called dilaton
in 10D supergravity-YM system, and the YM gauge coupling $g$
is determined by a vacuum expectation value of 10D dilaton
field $\phi_{10}$ as
\begin{eqnarray}
g &=& e^{\vev{\phi_{10}}/2}.
\nonumber
\end{eqnarray}
Furthermore, the 4D effective action (\ref{eq:seff})
is written in a so-called string frame, which is obtained
by Eq.~(\ref{eq:csugra}) with a dilatation gauge fixing,
\begin{eqnarray}
C|_0 &=& e^{-\phi_4} e^{K|_0/6},
\label{eq:stf}
\end{eqnarray}
where $\phi_4$ is a zero-mode of $\phi_{10}$,
which we call a 4D dilaton field.
The VEV of $\phi_4$ satisfies
\begin{eqnarray}
e^{-2 \vev{\phi_4}} &=&
e^{-2 \vev{\phi_{10}}} \prod_i {\cal A}^{(i)}
\ = \ g^{-4} \prod_i {\cal A}^{(i)}.
\label{eq:phi4vev}
\end{eqnarray}
The YM-field independent part of the K\"ahler potential $K$
in the action (\ref{eq:csugra}), that we denote $K^{(0)}$,
is well known as
\begin{eqnarray}
K^{(0)} &=&
\ln \frac{g^8}{2^7 \left(
\prod_i {\cal A}_i \right)^2
\prod_i {\rm Im}\, \tau_i},
\nonumber
\end{eqnarray}
determined from a coefficient of
Einstein-Hilbert term in dimensionally reduced
4D effective supergravity action.

In the string frame (\ref{eq:stf}), we expand the action
(\ref{eq:csugra}) in powers of YM-fields and compare the
corresponding terms in the power series to those in globally
supersymmetric effective action (\ref{eq:seff}).
In this way, we determine the K\"ahler potential $K$,
the superpotential $W$ and the gauge kinetic function $f_a$
in the 4D effective supergravity action~(\ref{eq:csugra}) as
\begin{eqnarray}
K &=& K^{(0)} + \sum_{i,j}\,\sum_{a,b}\,
\sum_{{\cal I}_{ab}} Z_{{\cal I}_{ab}}^{\bar{i}j}
{\rm tr} \left[ \bar\phi_{\bar{i}}^{{\cal I}_{ab}} e^{-V^a}
\phi_j^{{\cal I}_{ab}} e^{V^b} \right],
\label{eq:ksugra} \\
W &=&
\sum_{i,j,k}\,\sum_{a,b,c}\,
\sum_{{\cal I}_{ab},\,
{\cal I}_{bc},\,
{\cal I}_{ca}}
\lambda_{{\cal I}_{ab}
{\cal I}_{bc}
{\cal I}_{ca}}^{ijk}
{\rm tr} \left[
\phi_i^{{\cal I}_{ab}}
\phi_j^{{\cal I}_{bc}}
\phi_k^{{\cal I}_{ca}} \right],
\label{eq:wsugra} \\
f_a &=& \frac{1}{g^2} \prod_i {\cal A}^{(i)},
\label{eq:fsugra}
\end{eqnarray}
where
the K\"ahler metric
$Z_{{\cal I}_{ab}}^{\bar{i}j}$
and the holomorphic Yukawa couplings
$\lambda_{{\cal I}_{ab}
{\cal I}_{bc}
{\cal I}_{ca}}^{ijk}$
are given by
\begin{eqnarray}
Z_{{\cal I}_{ab}}^{\bar{i}j} &=&
e^{2 \vev{\phi_4}}
\tilde{Z}_{{\cal I}_{ab}}^{\bar{i}j},
\label{eq:kmsugra} \\
\lambda_{{\cal I}_{ab}
{\cal I}_{bc}
{\cal I}_{ca}}^{ijk}
&=& e^{3 \vev{\phi_4}} e^{-K^{(0)}/2}
\tilde\lambda_{{\cal I}_{ab}
{\cal I}_{bc}
{\cal I}_{ca}}^{ijk},
\label{eq:ycsugra}
\end{eqnarray}
and
$\tilde{Z}_{{\cal I}_{ab}}^{\bar{i}j}$,
$\tilde\lambda_{{\cal I}_{ab}
{\cal I}_{bc}{\cal I}_{ca}}^{ijk}$
are given in Eqs.~(\ref{eq:km}) and (\ref{eq:yc}),
respectively.
Note that the VEV of 4D dilaton field $\vev{\phi_4}$
in Eqs.~(\ref{eq:kmsugra}) and (\ref{eq:ycsugra})
is related to $g$ and ${\cal A}^{(i)}$
as shown in Eq.~(\ref{eq:phi4vev}).

\subsection{Moduli dependence}
\label{ssec:moduli}
As mentioned above, the YM gauge coupling constant $g$
is determined by a VEV of 10D dilaton field $\phi_{10}$
which is a dynamical scalar field.
Furthermore, the geometric (torus) parameters
$R_i$ and $\tau_i$ in the metric (\ref{eq:gtorus}) should
be considered as VEVs of dynamical fields originating from
10D metric $G_{MN}$. The zero-modes of these dynamical
fields are called moduli.
In the case of toroidal compactification, there exist
K\"ahler moduli and complex structure moduli whose VEVs
determine $R_i$ and $\tau_i$, respectively, in addition
to the zero-mode $\phi_4$ of 10D dilaton field $\phi_{10}$.

We assume (at least ${\cal N}=1$) supersymmetric vacuum
configurations (\ref{eq:vevs}) in this paper, and then
geometric moduli as well as $\phi_4$ form ${\cal N}=1$
supermultiplets described by chiral superfields.
Because these moduli fields are singlets under gauge
transformations of SYM sector, they do not feel magnetic
fluxes. Therefore, the formation of moduli supermultiplets
is the same one as a pure toroidal case without magnetic
fluxes, where we usually denote the dilaton, K\"ahler and
complex-structure moduli chiral superfields as
$S$, $T_i$ and $U_i$, respectively.
A suitable identification of
the VEVs of these fields are known~\cite{Cremades:2004wa} as
\begin{eqnarray}
{\rm Re}\,\vev{S}|_0 &\equiv& g^{-2} \prod_i {\cal A}^{(i)},
\qquad
{\rm Re}\,\vev{T_i}|_0 \ \equiv \ g^{-2} {\cal A}^{(i)},
\qquad
\vev{U_i}|_0 \ \equiv \ i \bar\tau_i.
\label{eq:stuvev}
\end{eqnarray}

An important task here is to identify
the dependence of 4D ${\cal N}=1$ effective supergravity
action (\ref{eq:csugra}) on the dilaton and moduli chiral
superfields $S$, $T_i$ and $U_i$. The K\"ahler and
superpotential as well as the gauge kinetic functions
in the action (\ref{eq:csugra}) are shown in
Eqs.~(\ref{eq:ksugra}), (\ref{eq:wsugra}) and
(\ref{eq:fsugra}), which are written in terms of VEVs
of moduli, namely, parameters $g$, $R_i$ and $\tau_i$.
Correct combinations of these parameters should be
promoted to moduli chiral superfields $S$, $T_i$ and $U_i$
following the relations (\ref{eq:stuvev})
up to a certain parameter-dependent rescaling
of YM superfields $V^a$ and $\phi_i^{{\cal I}_{ab}}$
in the action (\ref{eq:csugra}).
The rescaling is performed as
\begin{eqnarray}
V^a &\to& V^a, \qquad
\phi_i^{{\cal I}_{ab}} \ \to \
\alpha^{(i)}_{ab} \phi_i^{{\cal I}_{ab}},
\label{eq:rescale}
\end{eqnarray}
where,
for $\prod_r M^{(r)}_{ab} M^{(r)}_{bc} M^{(r)}_{ca} \ne 0$,
\begin{eqnarray}
\alpha^{(i)}_{ab} &=&
\frac{1}{g \sqrt{2\,{\rm Im}\,\tau_i}}\,
\left( \prod_r
\frac{{\cal A}^{(r)}}{\sqrt{2\,{\rm Im}\,\tau_r}} \right)^{1/2}
\exp \left[ -\sum_r \frac{\pi i}{{\rm Im}\,\tau_r}
\frac{\bar\zeta^{(r)}_{ab}}{M^{(r)}_{ab}}
{\rm Im}\,\zeta^{(r)}_{ab} \right]
\left( \frac{|M^{(i)}_{ab}|}{
\prod_{r \ne i} |M^{(r)}_{ab}|} \right)^{1/4},
\label{eq:fred}
\end{eqnarray}
so that all the parameters $g$, $R_i$ and $\tau_i$
can be promoted to superfields $S$, $T_i$ and $U_i$
through Eq.~(\ref{eq:stuvev})
with a proper holomorphicity in the superspace action.
In a case with some vanishing fluxes,
$\prod_r M^{(r)}_{ab} M^{(r)}_{bc} M^{(r)}_{ca}=0$,
the way of rescaling is shown in Appendix~\ref{sec:app}.

After the above rescaling and promotion
for $\prod_r M^{(r)}_{ab} M^{(r)}_{bc} M^{(r)}_{ca} \ne 0$,
we find moduli dependence of
the YM-field independent part of the
K\"ahler potential $K^{(0)}$,
the K\"ahler metric
$Z_{{\cal I}_{ab}}^{\bar{i}j}$,
the holomorphic Yukawa couplings
$\lambda_{{\cal I}_{ab}
{\cal I}_{bc}{\cal I}_{ca}}^{ijk}$
and the gauge kinetic functions $f_a$ as
\begin{eqnarray}
K^{(0)}(S,T,U) &=&
-\ln (S+\bar{S})
-\sum_r \ln(T_r + \bar{T}_r)
-\sum_r \ln(U_r + \bar{U}_r),
\label{eq:k0mod} \\
Z_{{\cal I}_{ab}}^{\bar{i}j}(S,T,U)
&=& \delta^{\bar{i}j}
\left( \frac{T_j+\bar{T}_{\bar{j}}}{2} \right)^{-1}
\left( \prod_r
\frac{U_r+\bar{U}_{\bar{r}}}{2} \right)^{-1/2}
\nonumber \\ && \times
\frac{1}{2^{5/2}}
\left( \frac{|M^{(j)}_{ab}|}
{\prod_{r \ne j} |M^{(r)}_{ab}|} \right)^{1/2}
\exp \left[ -\sum_r \frac{4 \pi}{U_r+\bar{U}_{\bar{r}}}
\frac{({\rm Im}\,\zeta^{(r)}_{ab})^2}{M^{(r)}_{ab}} \right],
\label{eq:kmmod} \\
\lambda_{{\cal I}_{ab}
{\cal I}_{bc}{\cal I}_{ca}}^{ijk}(U)
&=&
-\frac13\epsilon^{{\rm i}{\rm j}{\rm k}}
\delta_{{\rm i}}^{\ i}
\delta_{{\rm j}}^{\ j}
\delta_{{\rm k}}^{\ k}
\prod_r
\lambda^{(r)}_{
I^{(r)}_{ab}I^{(r)}_{bc}I^{(r)}_{ca}}(U),
\label{eq:ycmod} \\
f_a(S) &=& S.
\label{eq:famod}
\end{eqnarray}
For $a$, $b$ and $c$ satisfying
$M^{(r)}_{ab}M^{(r)}_{bc}M^{(r)}_{ca}>0$
(that is equivalent to
$M^{(r)}_{ac}M^{(r)}_{cb}M^{(r)}_{ba}<0$),
the function
$\lambda^{(r)}_{
I^{(r)}_{ab}I^{(r)}_{bc}I^{(r)}_{ca}}(U)$
in Eq.~(\ref{eq:ycmod}) is evaluated as
\begin{eqnarray}
\lambda^{(r)}_{
I^{(r)}_{ab}I^{(r)}_{bc}I^{(r)}_{ca}}(U) &=&
\left\{ \begin{array}{lll}
\lambda^{(r)}_{ab,c}(U)
&& (M^{(r)}_{ab}>0) \\*[5pt]
\lambda^{(r)}_{bc,a}(U)
&& (M^{(r)}_{bc}>0) \\*[5pt]
\lambda^{(r)}_{ca,b}(U)
&& (M^{(r)}_{ca}>0)
\end{array} \right.,
\label{eq:ycimod}
\end{eqnarray}
where
\begin{eqnarray}
\lambda^{(r)}_{ab,c}(U) &=&
\sum_{m=1}^{M^{(r)}_{ab}}
\delta_{I^{(r)}_{bc}+I^{(r)}_{ca}
-m M^{(r)}_{bc},\, I^{(r)}_{ab}}\,
\nonumber \\ && \times
\vartheta
\begin{bmatrix}
\frac{
M^{(r)}_{bc}I^{(r)}_{ca}
-M^{(r)}_{ca}I^{(r)}_{bc}
+m M^{(r)}_{bc}M^{(r)}_{ca}}{
M^{(r)}_{ab}M^{(r)}_{bc}M^{(r)}_{ca}}
\\ 0 \end{bmatrix} \left(
\bar\zeta^{(r)}_{ca}M^{(r)}_{bc}
-\bar\zeta^{(r)}_{bc}M^{(r)}_{ca},\,
iU_r M^{(r)}_{ab}M^{(r)}_{bc}M^{(r)}_{ca}
\right).
\nonumber
\end{eqnarray}
In a case with some vanishing fluxes,
$\prod_r M^{(r)}_{ab} M^{(r)}_{bc} M^{(r)}_{ca}=0$,
the expression of K\"ahler metric is shown
in Appendix~\ref{sec:app}.

Note that these functions of moduli
(\ref{eq:k0mod})-(\ref{eq:famod})
appear in the action (\ref{eq:csugra}) with
the K\"ahler potential (\ref{eq:ksugra}) and
the superpotential (\ref{eq:wsugra}) where
the YM fields $V^a$ and $\phi_i^{{\cal I}_{ab}}$
represent those after the rescaling (\ref{eq:fred}).
These results (\ref{eq:k0mod})-(\ref{eq:famod})
are consistent with those obtained in different
ways~\cite{Cremades:2004wa,DiVecchia:2008tm}.
A systematic formulation of 4D effective theory
respecting ${\cal N}=1$ superspace structure
presented here will be easily adopted to more
general systems of magnetized SYM theories
and D-branes.

\section{An example of model building}
\label{sec:model}
In this section, we indicate a possible direction of
phenomenological model building based on our formulation.
Starting from 10D $U(N)$ SYM theory with $N=8$, we assume
magnetic fluxes yielding $\tilde{N}=3$ and
$(N_1,N_2,N_3) \equiv (N_C,N_L,N_R)=(4,2,2)$
in Eq.~(\ref{eq:gsb}) that break YM symmetry
as $U(8) \to U(4)_C \times U(2)_L \times U(2)_R$.
We consider the case that this is further broken down to
$U(3)_C \times U(1)_{C'} \times U(2)_L
\times U(1)_{R'} \times U(1)_{R''}$
by Wilson-lines yielding
$\tilde{N}=5$ and $(N_1,N_2,N_3,N_4,N_5)
\equiv (N_C,N_{C'},N_L,N_{R'},N_{R''}) = (3,1,2,1,1)$.
The situation is realized~\cite{Ohki:2010nf}
by the following magnetic fluxes and Wilson-lines
for $r=1,2,3$:
\begin{eqnarray}
F_{2+2r,3+2r} &=& 2 \pi
\begin{pmatrix}
M^{(r)}_C {\bm 1}_4 & \ & \ \\
\ & M^{(r)}_L {\bm 1}_2 & \ \\
\ & \ & M^{(r)}_R  {\bm 1}_2
\end{pmatrix},
\nonumber \\
\zeta_r &=&
\begin{pmatrix}
\zeta^{(r)}_C {\bm 1}_3 & \ & \ & \ & \ \\
\ & \zeta^{(r)}_{C'} & \ & \ & \ \\
\ & \ & \zeta^{(r)}_L {\bm 1}_2 & \ & \ \\
\ & \ & \ & \zeta^{(r)}_{R'} & \ \\
\ & \ & \ & \ & \zeta^{(r)}_{R''}
\end{pmatrix},
\nonumber
\end{eqnarray}
where ${\bm 1}_N$ is a $N \times N$ unit matrix, and
all the nonvanishing entries take different values
from each other.

We embed the gauge symmetries $SU(3)_C$ and $SU(2)_L$
of the standard model into the above unbroken gauge groups
as $SU(3)_C \subset U(3)_C$ and $SU(2)_L \subset U(2)_L$.
Then, in order to obtain three generations of quarks and
leptons from the zero-mode degeneracy (\ref{eq:igen}) and
full-rank Yukawa matrices from the 10D gauge interaction,
the magnetic fluxes are determined, e.g., as
\begin{eqnarray}
(M^{(1)}_C,M^{(1)}_L,M^{(1)}_R) &=& (0,+3,-3),
\nonumber \\
(M^{(2)}_C,M^{(2)}_L,M^{(2)}_R) &=& (0,-1,0),
\nonumber \\
(M^{(3)}_C,M^{(3)}_L,M^{(3)}_R) &=& (0,0,+1),
\label{eq:model}
\end{eqnarray}
which correspond to
\begin{eqnarray}
M^{(1)}_C-M^{(1)}_L \ = \ -3, \qquad
M^{(1)}_L-M^{(1)}_R &=& +6, \qquad
M^{(1)}_R-M^{(1)}_C \ = \ -3, \nonumber \\
M^{(2)}_C-M^{(2)}_L \ = \ +1, \qquad
M^{(2)}_L-M^{(2)}_R &=& -1, \qquad
M^{(2)}_R-M^{(2)}_C \ = \ 0, \nonumber \\
M^{(3)}_C-M^{(3)}_L \ = \ 0, \qquad
M^{(3)}_L-M^{(3)}_R &=& -1, \qquad
M^{(3)}_R-M^{(3)}_C \ = \ +1.
\label{eq:f3gen}
\end{eqnarray}
In this case, supersymmetry conditions
(\ref{eq:abd}) and (\ref{eq:abfi}) are satisfied by
\begin{eqnarray}
{\cal A}^{(1)}/{\cal A}^{(2)}
={\cal A}^{(1)}/{\cal A}^{(3)}=3.
\label{eq:susy}
\end{eqnarray}

In this model, chiral superfields
$Q$, $U$, $D$, $L$, $N$, $E$, $H_u$ and $H_d$
carrying
the left-handed quark ($Q$),
the right-handed up-type quark ($U$),
the right-handed down-type quark ($D$),
the left-handed lepton ($L$),
the right-handed neutrino ($N$),
the right-handed electron ($E$),
the up-type Higgs particle ($H_u$) and
the down-type Higgs particle ($H_d$), respectively,
are found in $\phi_i^{{\cal I}_{ab}}$ as
\begin{eqnarray}
\phi_1^{{\cal I}_{ab}} &=&
\left(
\begin{array}{cc|c|cc}
\Omega_C^{(1)} & \Xi_{CC'}^{(1)} & 0 &
\Xi_{CR'}^{(1)} & \Xi_{CR''}^{(1)} \\
\Xi_{C'C}^{(1)} & \Omega_{C'}^{(1)} & 0 &
\Xi_{C'R'}^{(1)} & \Xi_{C'R''}^{(1)} \\
\hline
\Xi_{LC}^{(1)} & \Xi_{LC'}^{(1)} & \Omega_L^{(1)} &
H_u^K & H_d^K \\
\hline
0 & 0 & 0 & \Omega_{R'}^{(1)} & \Xi_{R'R''}^{(1)} \\
0 & 0 & 0 & \Xi_{R''R'}^{(1)} & \Omega_{R''}^{(1)}
\end{array}
\right),
\label{eq:phi1cont} \\
\phi_2^{{\cal I}_{ab}} &=&
\left(
\begin{array}{cc|c|cc}
\Omega_C^{(2)} & \Xi_{CC'}^{(2)} & Q^I & 0 & 0 \\
\Xi_{C'C}^{(2)} & \Omega_{C'}^{(2)} & L^I & 0 & 0 \\
\hline
0 & 0 & \Omega_L^{(2)} & 0 & 0 \\
\hline
0 & 0 & 0 & \Omega_{R'}^{(2)} & \Xi_{R'R''}^{(2)} \\
0 & 0 & 0 & \Xi_{R''R'}^{(2)} & \Omega_{R''}^{(2)}
\end{array}
\right),
\label{eq:phi2cont} \\
\phi_3^{{\cal I}_{ab}} &=&
\left(
\begin{array}{cc|c|cc}
\Omega_C^{(3)} & \Xi_{CC'}^{(3)} & 0 & 0 & 0 \\
\Xi_{C'C}^{(3)} & \Omega_{C'}^{(3)} & 0 & 0 & 0 \\
\hline
0 & 0 & \Omega_L^{(3)} & 0 & 0 \\
\hline
U^J & N^J & 0 & \Omega_{R'}^{(3)} & \Xi_{R'R''}^{(3)} \\
D^J & E^J & 0 & \Xi_{R''R'}^{(3)} & \Omega_{R''}^{(3)}
\end{array}
\right),
\label{eq:phi3cont}
\end{eqnarray}
where the rows and columns of matrices
correspond to $a=1,\ldots,5$ and $b=1,\ldots,5$,
respectively, and the indices $I,J=1,2,3$ and
$K=1,\ldots,6$ label generations.
Three generations of
$Q$, $U$, $D$, $L$, $N$, $E$ and
six generations of $H_u$ and $H_d$
are generated by the fluxes~(\ref{eq:model}).
The K\"ahler metric and holomorphic Yukawa couplings
for these superfields are easily derived from
Eqs.~(\ref{eq:kmmod}) and (\ref{eq:ycmod}).

Note that each of zero entries in the matrices
(\ref{eq:phi1cont}), (\ref{eq:phi2cont}) and
(\ref{eq:phi3cont}) represents eliminated components
due to the effect of chirality projection caused
by magnetic fluxes. However, because we require
some vanishing fluxes in Eq.~(\ref{eq:f3gen})
in order to obtain three generations of quarks and
leptons, there appear some massless exotic modes
$\Xi_{ab}^{(r)}$ as well as diagonal components
$\Omega_{a}^{(r)}$ (so-called open string moduli),
all of which feel zero fluxes.
Some of these modes can be eliminated
if we consider certain orbifold projections on
$r=2,3$ tori, that is,
a magnetized orbifold model~\cite{Abe:2008fi}.
More details of this model building and
phenomenological features at a low energy
will be reported in a separate paper~\cite{akos}.

\section{Conclusion}
\label{sec:conc}
We have presented 4D ${\cal N}=1$ superfield description
of 10D SYM theories compactified on magnetized tori which
preserve the ${\cal N}=1$ supersymmetry. Based on such a
description, we have derived 4D effective action for
massless zero-modes written in the ${\cal N}=1$ superspace.
We further identified moduli dependence of the effective
action by promoting the YM gauge coupling constant $g$ and
geometric parameters $R_i$ and $\tau_i$ to a dilaton,
K\"ahler and complex-structure moduli superfields.
The resulting effective supergravity action would be
useful for building phenomenological models
and for analyzing them systematically.

Although we have worked on 10D SYM theories in this paper,
it is straightforward to adopt our formulation to SYM in
lower-than-ten dimensional spacetime, in a similar way to
the one suggested in Ref.~\cite{ArkaniHamed:2001tb} without
magnetic fluxes. A local supersymmetry can be recovered
in 4D effective theories\footnote{
The corresponding supergravity actions in the original
higher-dimensional spacetime could be also written in
${\cal N}=1$ superspace as shown in the case of
five dimensions~\cite{Paccetti:2004ri}.}
following the procedure presented in Sec.~\ref{sec:sugra}.
Then, e.g., in type IIB orientifolds,
our formulation will be applied not only to magnetized D9
branes (a class of which is T-dual to intersecting D6 branes
in IIA side), but also to D5-D9~\cite{DiVecchia:2011mf} and
D3-D7 brane configurations with magnetic fluxes in extra dimensions.
Lower-dimensional brane configurations may allow the introduction of
supersymmetry-breaking branes sequestered from the visible sector,
which coincide with flavor structures in the visible sector
generated by magnetic fluxes.

The explicit moduli dependence of the superspace effective
action also allows us to study a moduli stabilization and a
supersymmetry breaking at a Minkowski minimum~\cite{Kachru:2003aw}
based on SYM theories, by minimizing the moduli and hidden-sector
potential generated by some combinations~\cite{Dudas:2006gr} of
nonperturbative effects and
a dynamical supersymmetry breaking~\cite{Intriligator:2006dd}.
Then, it would be possible to determine explicit forms of
soft terms in the visible sector generated by moduli-mediated
supersymmetry breaking (or mixed modulus-anomaly
mediation~\cite{Choi:2004sx}).
In such models, brane configurations in the higher-dimensional
spacetime might be detected by measuring supersymmetric flavor
structures at a low energy.
The formulation of 4D effective action presented here
would be suitable for such analyses.

\subsection*{Acknowledgement}
The work of H.~A. was supported by the Waseda University
Grant for Special Research Projects No.2011B-177.
The work of T.~K. is supported in part by a Grant-in-Aid for
Scientific Research No.~20540266 and the Grant-in-Aid for
the Global COE Program ``The Next Generation of Physics,
Spun from Universality and Emergence'' from the Ministry of
Education, Culture, Sports, Science and Technology of Japan.
The work of H.~O. is supported by the JSPS Grant-in-Aid for
Scientific Research (S) No. 22224003.
H.~A. and T.~K. thank the Yukawa Institute for Theoretical
Physics at Kyoto University.
Discussions during the YITP workshop~"Summer Institute 2011"
were useful to complete this work.

redpajama_set_name: RedPajamaArXiv


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text:
VIZ MEDIA ANNOUNCES HOME MEDIA RELEASE OF THE COMPLETE RANMA ½ ANIME SERIES IN NEW BLU-RAY STANDARD EDITION SETS
Beloved Martial Arts Anime Returns In Seven New Adventure-Packed Standard Edition Blu-ray Releases
San Francisco, CA, May 31, 2016 – VIZ Media, LLC (VIZ Media),a premiere company in the field of publishing, animation distribution, and global entertainment licensing, announces the release of new Blu-ray Standard Editions for all 7 RANMA ½ home media sets.
The RANMA ½ Blu-ray sets, which present the complete 161-episode series, are available now and sold individually with an MSRP of $49.99 U.S. / $57.99 CAN. Each set features 23 episodes across 3 discs, along with a variety of bonus material. Extras include convention highlights and RANMA ½ panel discussions, episode previews, series trailers, and a special 7-part "We Love Ranma" series that takes audiences behind the scenes, detailing the processes of the manga and anime release and the enduring popularity of this iconic action adventure series.
RANMA ½ Blu-ray episodes are presented in full 1080p High Definition resolution with the series' original, un-cropped 4:3 aspect ratio. Subtitled and English dubbed dialogue options are available for all episodes.
In the classic RANMA ½ anime series, Ranma Saotome never thought a martial arts training mission to China would turn out like this! An accidental dunk in a cursed spring changes boy Ranma into a young girl, every time he is splashed with cold water. Hot water reverses the effect - but only until next time!
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text:
» NGO
ROTA rolls out sixth batch of better connections programme
Reach Out to Asia (ROTA) has kicked off the sixth batch of the 'Better Connections' programme, which aims to improve the social wellbeing of migrant workers in Qatar by providing them with information and communication technology (ICT) equipment and skills.
On 23rd September, volunteers received in-depth training enabling them to offer technical support to companies in setting up computer rooms at migrant worker accommodation. Additionally, on 7th October, volunteers were taught how to provide workers with ICT literacy training to ensure they are equipped with the skills to utilise the facilities.
The 'Better Connections' programme was officially launched by the Ministry of Transport and Communications, with ROTA's role being sponsored by Msheireb Properties, a subsidiary of Qatar Foundation for Education, Science and Community Development (QF). Msheireb Properties' mission is to change the way people think about urban living through innovations that encourage social interaction, a respect for culture, and greater care for the environment. Microsoft Qatar and the Sheikh Thani Bin Abdullah Foundation for Humanitarian Services (RAF) are implementing partners in the programme.
Waleed al-Amri, a 'Better Connections' volunteer, said: "ROTA's Better Connections training programme provides volunteers with the opportunity to learn new skills, and aims to improve the lives of migrant workers in Qatar, allowing for better communication with their families and communities back home." ROTA, a member of Qatar Foundation, joined 'Better Connections' in 2015. Since then, the programme has trained a total of 148 volunteers who have gone on to train more than 745 'Digital Champions' as part of the social development initiative.
Anwar Abdulbaki, ICT for Education Manager at ROTA, said: "We are extremely thankful to the enthusiastic group of volunteers who are fully committed to improving the social well-being of migrant workers by giving them the opportunity to gain essential ICT skills. Through the Better Connections programme, our volunteers and partner companies are actively contributing to the Qatar National Vision 2030 and to building a knowledge-based economy in Qatar by investing in human capital."
Noora Al Rumaihi, Supervisor in Communication and Public Relations at Msheireb Properties, said: "Msheireb Properties is committed to giving back to communities in Qatar. Through our continued partnership with ROTA and programmes such as Better Connections, we are more effectively able to continue to enact solutions to local challenges and to facilitate significant differences in the lives of others while also creating a long-lasting impact in the communities where they live and work."
Reach Out To Asia
Better Connections
Msheireb Properties
Waleed al-Amri
MOTC to start trial for Automatic Rapid Transit system
QCDC wraps up successful first edition of Career Academy
From NGO
Qatar's charities bring sustainability to the fore
HE Sheikha Hind participates in UN discussion on driving sustainable development through...
QNRF announces 54 winners in the 24th cycle of its Undergraduate Research Experience...
QF and Wanda Group give students from Qatar the chance to star at FIFA Women's France...
DIFI attends 73rd session of the Committee on the Elimination of Discrimination against...


text:
https://apnews.com/article/north-america-ap-top-news-sexual-assault-sexual-misconduct-movies-7a400aa4694b404190a6882096d70d67
CA State Wire
Los Angeles prosecutors reviewing Spacey sex assault case
April 11, 2018 GMT
FILE - In this Oct. 27, 2017 file photo, Kevin Spacey presents the award for excellence in television at the BAFTA Los Angeles Britannia Awards in Beverly Hills, Calif. Los Angeles County prosecutors say they are reviewing a sexual assault case against Kevin Spacey. The district attorney's office spokesman Greg Risling said Wednesday, April 11, that sheriff's investigators presented the case to prosecutors April 5. No further details were provided. He is also under police investigation for sexual assault in London, according to British media. (Photo by Chris Pizzello/Invision/AP, File)
LOS ANGELES (AP) — Los Angeles County prosecutors are reviewing sexual assault allegations against Kevin Spacey for possible criminal charges, authorities said Wednesday.
Sheriff's investigators presented the allegations on April 5 to the district attorney's task force on sex crimes in the entertainment industry, a sheriff's statement said.
District attorney's spokesman Greg Risling confirmed that he case was under consideration by prosecutors.
A man alleges that Spacey sexually assaulted him in West Hollywood in October of 1992, the sheriff's statement said.
Investigators from the department's Special Victims Bureau began an investigation in December, a few weeks after the actor was first publicly accused of sexual misconduct by actor Anthony Rapp.
No further details about the investigation were provided.
Spacey's representative Laura Johnson declined comment in an email.
British media have reported that Spacey, 58, is also under police investigation for sexual assault in London.
The Academy Award winner was first publicly accused of sexual misconduct in October. He quickly lost roles in the film "All the Money in the World" and the TV show "House of Cards," and future projects were shelved.
Spacey apologized to Rapp and his former publicist said he was seeking unspecified treatment.


text:
Top 5 Current NFL Players
Luis Cardenas, Contributor
Number 1: Lamar Jackson (QB)
Lamar Demeatrice Jackson Jr. is an American football quarterback for the Baltimore Ravens. In 2019, he was the NFL's Most Valuable Player, the second player to receive a unanimous selection for the award, and is the fourth African-American quarterback to be named MVP. He was born on January 7, 1997.
Number 2: Russell Wilson (QB)
Russell Carrington Wilson is an American football quarterback for the Seattle Seahawks. Wilson initially played football and baseball for North Carolina State University from 2008 to 2010 before transferring to Wisconsin. He was born on November 29, 1988.
Number 3: Aaron Donald (DT)
Aaron Charles Donald is an American football defensive tackle for the Los Angeles Rams. He played college football at Pittsburgh, where he was recognized as a unanimous All-American. He was drafted by the Rams with the thirteenth pick in the first round of the 2014 NFL Draft. He was born on May 23, 1991.
Number 4: Patrick Mahomes ll (QB)
Patrick Lavon Mahomes ll is an American football quarterback for the Kansas City Chiefs. He is the son of former Major League Baseball pitcher Pat Mahomes. He initially played college football and college baseball at Texas Tech University, He was born on September 17, 1995.
Number 5: Michael Thomas (WR)
Michael William Thomas Jr. is an American football wide receiver for the New Orleans Saints. He played college football at Ohio State and was born on March 3, 1993. Thomas holds the record for most receptions by a player through his first four seasons with 470 along with the most receptions by a player in a single season with 149.
My name is Luis Cardenas, and I am 16 years old. I am from Oxnard, California, and have always lived in Oxnard. I live with my mom, dad, and my two sisters....
English Premier League: The Most Watched League in Professional Soccer
HHS Homecoming Game
Hueneme Football Team in the Playoffs!
Los Angeles Lakers Win NBA Finals
What it is Like to be a Cheerleader
Fans Went Crazy Because the Lakers Won
Two Baseball Games are Postponed Because of Coronavirus Cases
Nuggets vs Clippers


text:
Gods Good Man
Marie Corelli
God's Good Man, by Marie Corelli
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GOD'S GOOD MAN
A Simple Love Story
By MARIE CORELLI
AUTHOR OF "THE TREASURE OF HEAVEN," "THELMA," "A ROMANCE OF TWO WORLDS," "THE MASTER CHRISTIAN," ETC.
THE LIVING ORIGINAL
"THE REVEREND JOHN WALDEN"
AND HIS WIFE
THIS SIMPLE LOVE STORY
AFFECTIONATELY DEDICATED
"THERE WAS A MAN SENT FROM GOD WHOSE NAME WAS JOHN." NEW TESTAMENT
It was May-time in England.
The last breath of a long winter had blown its final farewell across the hills,--the last frost had melted from the broad, low-lying fields, relaxing its iron grip from the clods of rich, red-brown earth which, now, soft and broken, were sprouting thick with the young corn's tender green. It had been a hard, inclement season. Many a time, since February onward, had the too-eagerly pushing buds of trees and shrubs been nipped by cruel cold,--many a biting east wind had withered the first pale green leaves of the lilac and the hawthorn,--and the stormy caprices of a chill northern. Spring had played havoc with all the dainty woodland blossoms that should, according to the ancient 'Shepherd's Calendar' have been flowering fully with the daffodils and primroses. But during the closing days of April a sudden grateful warmth had set in,--Nature, the divine goddess, seemed to awaken from long slumber and stretch out her arms with a happy smile,--and when May morning dawned on the world, it came as a vision of glory, robed in clear sunshine and girdled with bluest skies. Birds broke into enraptured song,--young almond and apple boughs quivered almost visibly every moment into pink and white bloom,--cowslips and bluebells raised their heads from mossy corners in the grass, and expressed their innocent thoughts in sweetest odour--and in and through all things the glorious thrill, the mysterious joy of renewed life, hope and love pulsated from the Creator to His responsive creation.
It was May-time;--a real 'old-fashioned' English May, such as Spenser and Herrick sang of:
"When all is yclad With blossoms; the ground with grass, the woodes With greene leaves; the bushes with blossoming buddes,"
and when whatever promise our existence yet holds for us, seems far enough away to inspire ambition, yet close enough to encourage fair dreams of fulfilment. To experience this glamour and witchery of the flowering-time of the year, one must, perforce, be in the country. For in the towns, the breath of Spring is foetid and feverish,--it arouses sick longings and weary regrets, but scarcely any positive ecstasy. The close, stuffy streets, the swarming people, the high buildings and stacks of chimneys which only permit the narrowest patches of sky to be visible, the
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