Cosmic Strings and Cosmic Superstrings

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  • Cosmic Strings and Cosmic Superstrings

    Mairi Sakellariadou a

    aDepartment of Physics, Kings College, University of London, Strand, London WC2R 2LS, U.K.

    In these lectures, I review the current status of cosmic strings and cosmic superstrings. I rst discuss topo-

    logical defects in the context of Grand Unied Theories, focusing in particular in cosmic strings arising as gauge

    theory solitons. I discuss the reconciliation between cosmic strings and cosmological ination, I review cosmic

    string dynamics, cosmic string thermodynamics and cosmic string gravity, which leads to a number of interesting

    observational signatures. I then proceed with the notion of cosmic superstrings arising at the end of brane ina-

    tion, within the context of brane-world cosmological models inspired from string theory. I discuss the dierences

    between cosmic superstrings and their solitonic analogues, I review our current understanding about the evolution

    of cosmic superstring networks, and I then briey describe the variety of observational consequences, which may

    help us to get an insight into the stringy description of our Universe.

    1. Introduction

    Provided our understanding about unicationof forces and big bang cosmology are correct,it is natural to expect that topological defects,appearing as solutions to many particle physicsmodels of matter, could have formed naturallyduring phase transitions followed by sponta-neously broken symmetries, in the early stagesof the evolution of the Universe. Certain types oftopological defects (local monopoles and local do-main walls) may lead to disastrous consequencesfor cosmology, hence being undesired, while oth-ers (cosmic strings) may play a useful role.

    Cosmic strings [1] are linear topological defects,analogous to ux tubes in type-II superconduc-tors, or to vortex laments in superuid helium.These objects gained a lot of interest in the 1980sand early 1990s, since they oered a potential al-ternative to the cosmological ination for the ori-gin of initial density uctuations leading to theCosmic Microwave Background (CMB) temper-ature anisotropies and the observed structure inthe Universe. They however lost their appeal,when it was found that they lead to inconsisten-cies in the power spectrum of the CMB. It waslater shown [2] that cosmic strings are genericallyformed at the end of an inationary era, withinthe framework of Supersymmetric Grand Unied

    Theories (SUSY GUTs). Hence cosmic stringshave to be included as a sub-dominant partnerof ination. This theoretical support gave a newboost to the eld of cosmic strings, a boost whichhas been more recently enhanced when it wasshown that cosmic superstrings [3] (fundamentalor one-dimensional Dirichlet branes) can play therole of cosmic strings, in the framework of brane-world cosmologies.

    A realistic cosmological scenario necessitatesthe input of high energy physics; any models de-scribing the early stages of the evolution of theUniverse have their foundations in general rela-tivity and high energy physics. Comparing thetheoretical predictions of such models against cur-rent astrophysical and cosmological data, resultsto either their acceptance or their rejection, whilein the rst case it also xes the free parametersof the models (see e.g., Ref. [4,5]). In particu-lar, by studying the properties of cosmic super-string networks and comparing their phenomeno-logical consequences against observational data,we expect to pin down the successful and natu-ral inationary model and get some insight intothe stringy description of the Universe. Cosmicstrings/superstrings represent a beautiful exam-ple of the strong and fruitful link between cos-mology and high energy physics.

    In what follows, I will summarise the material I

    Nuclear Physics B (Proc. Suppl.) 192193 (2009) 6890

    0920-5632/$ see front matter 2009 Elsevier B.V. All rights reserved.


  • had presented in my lectures during the summerschool at Carge`se (June 2008) 1. I will highlightonly certain aspects of the subject, which I con-sider either more important due to their obser-vational consequences, or more recently obtainedresults.

    2. Topological Defects in GUTs

    In the framework of the hot big bang cosmolog-ical model, the Universe was originally at a veryhigh temperature, hence the initial equilibriumvalue of the Higgs eld , which plays the roleof the order parameter, was at = 0. Since thePlanck time, the Universe has, through its expan-sion, steadily cooled down and a series of phasetransitions followed by Spontaneously SymmetryBreaking 2 (SSBs) took place in the frameworkof GUTs. Such SSBs may have left behind topo-logical defects as false vacuum remnants, via theKibble mechanism [6].

    The formation or not of topological defectsand the determination of their type, depend onthe topology of the vacuum manifold Mn. Theproperties of Mn are described by the kth ho-motopy group k(Mn), which classies distinctmappings from the k-dimensional sphere Sk intothe manifoldMn. Consider the symmetry break-ing of a group G down to a subgroup H of G.If Mn = G/H has disconnected components equivalently, if the order k of the non-trivial ho-motopy group is k = 0 two-dimensional de-fects, called domain walls, form. The space-timedimension d of the defects is given in terms ofthe order of the non-trivial homotopy group byd = 4 1 k. If Mn is not simply connected equivalently, if Mn contains loops which can-not be continuously shrunk into a point cosmicstrings form. A necessary, but not sucient, con-dition for the existence of stable strings is thatthe fundamental group 1 of Mn, is non-trivial,or multiply connected. Cosmic strings are linear-like defects, d = 2. If Mn contains unshrinkablesurfaces, then monopoles form; k = 1, d = 1. IfMn contains non-contractible three-spheres, then1 concept of spontaneous symmetry breaking has itsorigin in condensed matter physics.

    event-like defects, textures, form; k = 3, d = 0.Depending on whether the symmetry is local

    (gauged) or global (rigid), topological defects arerespectively, local or global. The energy of localdefects is strongly conned, while the gradientenergy of global defects is spread out over thecausal horizon at defect formation. Global de-fects having long range density elds and forces,can decay through long-range interactions, hencethey do not contradict observations, while localdefects may be undesirable for cosmology. Inwhat follows, I will discuss local defects, since weare interested in gauge theories, being the morephysical ones 3. Patterns of symmetry breakingwhich lead to the formation of local monopolesor local domain walls are ruled out, since theyshould soon dominate the energy density of theUniverse and close it, unless an inationary eratook place after their formation. This is one ofthe reasons for which cosmological ination aperiod in the earliest stages of the evolution ofthe Universe, during which the Universe could bein an unstable vacuum-like state having high en-ergy density, which remained almost constant was proposed. Local textures are insignicant incosmology since their relative contribution to theenergy density of the Universe decreases rapidlywith time.

    Even in the absence of a non-trivial topologyin a eld theory, it may still be possible to havedefect-like solutions, since defects may be em-bedded in such topologically trivial eld theories.However, while stability of topological defects isguaranteed by topology, embedded defects are ingeneral unstable under small perturbations.

    Let me discuss the genericity of cosmic stringformation in the context of SUSY GUTs, whichcontain a large number of SSB patterns leadingfrom a large gauge group GGUT to the Stan-dard Model (SM) gauge group GSM SU(3)CSU(2)L U(1)Y. The minimum rank of GGUThas to be at least equal to 4, to contain the GSMas a subgroup; we set the upper bound on therank r of the group to be r 8. The embeddingsof GSM in GGUT must be such that there is an

    3Note that when we say cosmic strings we refer to localone-dimensional topological defects.

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 6890 69

  • agreement with the SM phenomenology and es-pecially with the hypercharges of the known par-ticles. The large gauge group GGUT must includea complex representation, needed to describe theSM fermions, and it must be anomaly free. A de-tailed investigation [2] has concluded that GGUTcould be either one of SO(10), E6, SO(14), SU(8),SU(9); ipped SU(5) and [SU(3)]3 are includedwithin this list as subgroups of SO(10) and E6,respectively. The formation of domain walls ormonopoles, necessitates an era of supersymmet-ric hybrid ination to dilute them. ConsideringGUTs based on simple gauge groups, the type ofsupersymmetric hybrid ination will be of the F-type. The baryogenesis mechanism will be ob-tained via leptogenesis, either thermal or non-thermal leptogenesis. Finally, to ensure the sta-bility of proton, the discrete symmetry Z2, whichis contained in U(1)BL, must be kept unbrokendown to low energies; the successful SSB schemesshould end at GSM Z2. Taking all these con-siderations into account, a detailed study of allSSB schemes leading from a GGUT down to theGSM, by one or more intermediate steps, showsthat cosmic strings are generically formed at theend of hybrid ination.

    The results [2] can be summarised as follows: Ifthe large gauge group GGUT is the SO(10), thencosmic strings formation is unavoidable. Thegenericity of string formation in the case thatthe large gauge group is the E6, depends uponwhether one considers thermal or non-thermalleptogenesis. More precisely, for non-thermalleptogenesis, cosmic string formation is unavoid-able, while for thermal leptogenesis, cosmic stringformation arises in 98% of the acceptable SSBschemes. If the requirement of having Z2 un-broken down to low energies is relaxed and ther-mal leptogenesis is considered as being the mech-anism for baryogenesis, then cosmic string for-mation accompanies hybrid ination in 80% ofthe SSB schemes. The SSB schemes of eitherSU(6) or SU(7), as the large gauge group, downto the GSM, which could accommodate an ina-tionary era with no defect (of any kind) at latertimes are inconsistent with proton lifetime mea-surements, while minimal SU(6) and SU(7) donot predict neutrino masses, implying that these

    models are incompatible with high energy physicsphenomenology. Higher rank groups, namelySO(14), SU(8) and SU(9), should in general leadto cosmic string formation at the end of hybridination. In all these schemes, cosmic string for-mation is sometimes accompanied by the forma-tion of embedded strings. The strings which format the end of hybrid ination have a mass whichis proportional to the inationary scale.

    3. Cosmic Strings and Ination

    An appealing solution to the drawbacks of thestandard hot big bang model is to introduce,during the very early stages of the evolution ofthe Universe, a period of accelerated expansion,known as cosmological ination [7]. The ina-tionary era took place when the Universe wasin an unstable vacuum-like state at a high en-ergy density, leading to a quasi-exponential ex-pansion. The combination of the hot big bangmodel and the inationary scenario provides atpresent the most comprehensive picture of theUniverse at our disposal. Ination ends when theHubble parameter H =

    8/(3M2Pl) (where

    denotes the energy density and MPl stands forthe Planck mass) starts decreasing rapidly. Theenergy stored in the vacuum-like state gets trans-formed into thermal energy, heating up the Uni-verse and leading to the beginning of the standardhot big bang radiation-dominated era.

    Ination is based on the basic principles of gen-eral relativity and eld theory, while when theprinciples of quantum mechanics are also con-sidered, it provides a successful explanation forthe origin of the large scale structure, associatedwith the measured temperature anisotropies inthe CMB spectrum. Despite its remarkable suc-cess, ination still remains a paradigm in searchof model. An inationary model should be in-spired from a fundamental theory, while its pre-dictions should be tested against current data. Inaddition, releasing the present Universe form itsacute dependence on the initial data, ination isfaced with the challenging task of proving itselfgeneric [8], in the sense that ination would takeplace without ne-tuning of the initial conditions.

    Theoretically motivated inationary models

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  • can be built in the context of supersymmetryor Supergravity (SUGRA). N=1 supersymmetrymodels contain complex scalar elds which of-ten have at directions in their potential, thusoering natural candidates for inationary mod-els. In this framework, hybrid ination driven byF-terms or D-terms is the standard inationarymodel, leading generically to cosmic string for-mation at the end of ination. Hybrid ination isbased on Einsteins gravity but is driven by thefalse vacuum. The inaton eld rolls down itspotential while another scalar eld is trapped inan unstable false vacuum. Once the inaton eldbecomes much smaller than some critical value, aphase transition to the true vacuum takes placeand ination ends. F-term ination is potentiallyplagued with the Hubble-induced mass problem4 (-problem), while D-term ination avoids it.

    F-term ination can be naturally accommo-dated in the framework of GUTs, when a GGUTis broken down to the GSM, at an energy scaleMGUT according to the scheme


    +H2GSM ,

    where +, is a pair of GUT Higgs super-elds in non-trivial complex conjugate represen-tations, which lower the rank of the group by oneunit when acquiring non-zero vacuum expectationvalue. The inationary phase takes place at the

    beginning of the symmetry breaking H1Minfl H2.

    The gauge symmetry is spontaneously broken byadding F-terms to the superpotential. The Higgsmechanism leads generically [2] to Abrikosov-Nielsen-Olesen strings, called F-term strings.

    F-term ination is based on the globally super-symmetric renormalisable superpotential

    WFin = S(+ M2) , (1)where S is a GUT gauge singlet left handed su-pereld and , M are two constants (M has di-mensions of mass) which can be taken positivewith eld redenition.

    4In supergravity theories, the supersymmetry breaking istransmitted to all elds by gravity, and thus any scalareld, including the inaton, gets an eective mass of theorder of the expansion rate H during ination.

    The scalar potential, as a function of the scalarcomplex component of the respective chiral su-perelds , S, is

    V (+, , S) = |F+ |2 + |F |2 + |FS |2




    g2aD2a . (2)

    The F-term is such that Fi |W/i|=0,where we take the scalar component of the super-elds once we dierentiate with respect to i =, S. The D-terms are Da = i (Ta)


    j + a,with a the label of the gauge group generators Ta,ga the gauge coupling, and a the Fayet-Iliopoulosterm. By denition, in the F-term ination thereal constant a is zero; it can only be nonzero ifTa generates an extra U(1) group. In the contextof F-term hybrid ination the F-terms give riseto the inationary potential energy density whilethe D-terms are at along the inationary trajec-tory, thus one may neglect them during ination.

    The potential, has one valley of local minima,V = 2M4, for S > M with + = = 0, andone global supersymmetric minimum, V = 0, atS = 0 and + = = M . Imposing initiallyS M , the elds quickly settle down the valleyof local minima. Since in the slow-roll ination-ary valley the ground state of the scalar poten-tial is non-zero, supersymmetry is broken. In thetree level, along the inationary valley the poten-tial is constant, therefore perfectly at. A slopealong the potential can be generated by includingone-loop radiative corrections. Hence, the scalarpotential gets a little tilt which helps the inatoneld S to slowly roll down the valley of minima.The one-loop radiative corrections to the scalarpotential along the inationary valley lead to theeective potential [4]

    V Fe(|S|) = 2M4{1 +


    [2 ln


    + (z + 1)2 ln(1 + z1)

    + (z 1)2 ln(1 z1)]}


    with z = |S|2/M2, and N stands for the dimen-sionality of the representation to which the com-plex scalar components +, of the chiral su-perelds +, belong. This implies that the

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 6890 71

  • eective potential, Eq. (3), depends on the par-ticular symmetry breaking scheme considered.

    D-term ination can be easily implementedwithin high energy physics (e.g., SUSY GUTs,SUGRA, or string theories) and it avoids the-problem. Within D-term ination, the gaugesymmetry is spontaneously broken by introduc-ing Fayet-Iliopoulos (FI) D-terms. In standardD-term ination, the constant FI term gets com-pensated by a single complex scalar eld at theend of the inationary era, which implies thatstandard D-term ination ends always with theformation of cosmic strings, called D-term strings.A supersymmetric description of the standard D-term ination is insucient, since the inatoneld reaches values of the order of the Planckmass, or above it, even if one concentrates onlyaround the last 60 e-folds of ination. Thus, D-term ination has to be studied in the context ofsupergravity [4,5]).

    Standard D-term ination requires a scheme

    GGUT U(1) MGUT HU(1) Min+

    H GSM .

    It is based on the superpotential

    W = S+ , (4)

    where S,+, are three chiral superelds and is the superpotential coupling. It assumesan invariance under an Abelian gauge groupU(1), under which the superelds S,+,have charges 0, +1 and 1, respectively. It alsoassumes the existence of a constant FI term .In D-term ination the superpotential vanishesat the unstable de Sitter vacuum (anywhere elsethe superpotential is non-zero), implying thatwhen the superpotential vanishes, D-term ina-tion must be studied within a non-singular for-mulation of supergravity. Various formulationsof eective supergravity can be constructed fromthe superconformal eld theory. To constructa formulation of supergravity with constant FIterms from superconformal theory, one nds [9]that under U(1) gauge transformations in the di-rections in which there are constant FI terms ,the superpotential W must transform as W =i

    iW = i(g/M2Pl)W ; one cannot keep any

    longer the same charge assignments as in stan-dard supergravity.

    D-term inationary models can be built withdierent choices of the Kahler geometry. Variouscases have been explored in the literature. Thesimplest case is that of D-term ination withinminimal supergravity [4]. It is based on

    Kmin =


    |i|2 = ||2 + |+|2 + |S|2 , (5)

    with fab(i) = ab.Another example is that of D-term ination

    based on Kahler geometry with shift symmetry,

    Kshift =1

    2(S + S)2 + |+|2 + ||2 , (6)

    and minimal structure for the kinetic function [5].One can also consider consider [5] a Kahler po-

    tential with non-renormalisable terms:

    Knonrenorm = |S|2 + |+|2 + ||2


    ( |S|2M2Pl



    ( |S|2M2Pl

    )||2 + b |S|


    M2Pl, (7)

    where f are arbitrary functions of (|S|2/M2Pl)and the superpotential is given in Eq. (4).

    Having the superpotential, one must proceedin the same way as in F-term ination and writedown the three level scalar potential and then in-clude the one-loop radiate corrections.

    Let me nally note that dierent ap-proaches [10] have been proposed in order toavoid cosmic string formation in the context of D-term ination. For example, one can add a non-renormalisable term in the potential, or add anadditional discrete symmetry, or consider GUTmodels based on non-simple groups, or nally in-troduce a new pair of charged superelds so thatcosmic string formation is avoided within D-termination.

    4. Cosmic String Dynamics

    The world history of a cosmic string can be ex-pressed by a two-dimensional surface in the four-dimensional string world-sheet:

    x = x(a) , a = 0, 1 ; (8)

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  • the world-sheet coordinates 0, 1 are arbitraryparameters, 0 is time-like and 1 ( ) is space-like.

    Over distances that are large compared to thewidth of the string, but small compared to thehorizon size, solitonic cosmic strings can be con-sidered as one-dimensional objects and their mo-tion can be well-described by the Nambu-Gotoaction. Thus, the string equation of motion, inthe limit of a zero thickness string, is derived fromthe Goto-Nambu eective action

    S0[x] =

    d2 , (9)where = det(ab) with ab = gx


    ,b and

    stands for the linear mass density, with T 2c ,where Tc is the critical temperature of the phasetransition followed by SSB leading to cosmicstring formation. By varying the action, Eq. (9),with respect to x(a), and using d = abdab,we get the string equation of motion:

    x ;a,a +

    abx,ax,b = 0 ; (10)

    is the four-dimensional Christoel symbol.We have neglected the friction [11], due to the

    scattering of thermal particles o the string. Forstrings formed at the grand unication scale, fric-tion is important only for a very short period oftime. For strings formed at a later phase transi-tion (e.g., closer to the electroweak scale), frictionwould dominate their dynamics through most ofthe thermal history of the Universe.

    By varying the action with respect to the met-ric, the string energy-momentum tensor reads

    T g =

    d2abx,ax,b(4)(xx(a)) .

    In an expanding Universe, the cosmic stringequation of motion is most conveniently writtenin comoving coordinates, where the Friedmann-Lematre-Robertson-Walker (FLRW) metrictakes the form ds2 = a2()[d2 dr2] ; a() isthe cosmic scale factor in terms of conformal time (related to cosmological time t, by dt = ad).Under the gauge condition 0 = , the comovingspatial string coordinates, x(, ), are writtenas a function of , and the length parameter .

    For a string moving in a FLRW Universe, theequation of motion, Eq. (10), can be simpliedin the gauge for which the unphysical parallelcomponents of the velocity vanish,

    x x = 0 ; (11)overdots and primes denote derivatives with re-spect to and , respectively. In these coordi-nates, the Goto-Nambu action yields the follow-ing string equation of motion in a FLRW metric:

    x+ 2



    )x(1 x2) =



    ). (12)

    The string energy per unit , in comoving units,is

    x2/(1 x2), implying that the string

    energy is ad. One usually xes entirely the

    gauge by choosing so that = 1 initially.

    The string equation of motion is much simplerin Minkowski space-time. Equation (10) for atspace-time simplies to

    a(abx,b) = 0 . (13)

    We impose the conformal gauge

    x x = 0 , x2 + x2 = 0 ; (14)overdots and primes denote derivatives with re-spect to 0 and 1, respectively. In this gaugethe string equation of motion is just a two-dimensional wave equation:

    x x = 0 . (15)To x entirely the gauge, we set t x0 = 0 ,which allows us to write the string trajectoryas the three-dimensional vector x(, t), where1 , the space-like parameter along the string.Hence, the constraint equations, Eq. (14), and thestring equation of motion, Eq. (15), become

    x x = 0 , x2 + x2 = 1 , x x = 0 . (16)The above equations imply that the string movesperpendicularly to itself with velocity x, that is proportional to the string energy, and thatthe string acceleration in the string rest frameis inversely proportional to the local string cur-vature radius. A curved string segment tends tostraighten itself, resulting to string oscillations.

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 6890 73

  • The general solution to the string equation ofmotion in at space-time, Eq. (16c), is

    x =1

    2[a( t) + b( + t)] , (17)

    where a( t) and b( + t) are two continuousarbitrary functions which satisfy

    a2 = b2 = 1 . (18)

    Hence, is the length parameter along the three-dimensional curves a(),b().

    The Goto-Nambu action describes to a goodapproximation cosmic string segments which areseparated. However, it leaves unanswered the is-sue of what happens when strings cross; a studywhich necessitates full eld theory. When twostrings of the same type collide, they may eitherpass simply through one another, or they mayreconnect (intercommute). A necessary, but notsucient, condition for string reconnection is thatthe initial and nal congurations be kinemati-cally allowed in the innitely thing string approx-imation. Numerical simulations (and analyticalestimates) of type-II (and weakly type-I) stringsin the Abelian Higgs model suggest that the prob-ability that a pair of strings will reconnect, afterthey intersect, is close to unity. The results arebased on lattice simulations of the correspondingclassical eld congurations in the Abelian Higgsmodel; the internal structure of strings is highlynon-linear, and thus dicult to treat via analyt-ical means. String-string and self-string intersec-tions lead to the formation of new long strings andloops. String intercommutations produce discon-tinuities, kinks, in x and x on the new stringsegments at the intersection point, composed ofright- and left-moving pieces travelling along thestring at the speed of light.

    The rst analytical studies of the evolution ofa cosmic string network have shown [12] the ex-istence of scaling, in the sense that the stringnetwork can be characterised by a single lengthscale, roughly the persistence length or the inter-string distance which grows with the horizon.This important property of cosmic strings ren-ders them cosmologically acceptable, in contrastto local monopoles or domain walls. Early nu-merical simulations have shown [13] that the typ-

    ical curvature radius of long strings and the char-acteristic distance between the strings are bothcomparable to the evolution time t. The energydensity of super-horizon 5 strings in the scal-ing regime is given (in the radiation-dominatedera) by long = t

    2 , where is a numeri-cal coecient ( = 20 10). Assuming that thesuper-horizon strings are characterised by a singlelength scale (t), implies

    (t) = 1/2t . (19)

    The typical distance between the nearest stringsegments and the typical curvature radius of thestrings are both of the order of . These resultsagree with the picture of the scale-invariant evo-lution of the string network and with the one-scale hypothesis. Further numerical investiga-tions however revealed dynamical processes, suchas the production of small sub-horizon loops, atscales much smaller than [14]. In responseto these ndings, a three-scale model was devel-oped [15] which describes the network in thermsof three scales: the energy density scale , a corre-lation length along the string, and a scale re-lated to local structure on the string. The small-scale structure (wiggliness), which oers an expla-nation for the formation of the small sub-horizonsized loops, is basically developed through inter-sections of long string segments. Aspects of thethree-scale model have been checked [16] evolvinga cosmic string network in Minkowski space-time.

    The sub-horizon strings (loops), their size dis-tribution, and the mechanism of their formationremained for years the least understood parts ofthe string evolution. Recently, numerical sim-ulations of cosmic string evolution in a FLRWUniverse (see, Fig. 1), found [17] evidence of ascaling regime for the cosmic string loops in theradiation- and matter-dominated eras down tothe hundredth of the horizon time. The scal-ing was found without considering any gravita-tional back reaction eect; it was just the result

    5Often in the literature, strings are divided into twoclasses: string loops and innite, or long, strings. How-ever, in numerical simulations strings are always loops, inthe sense that they do not have open ends. Hence, theterm loops corresponds to sub-horizon string loops, whilethe term innite strings corresponds to super-horizonstring loops.

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 689074

  • Figure 1. Snapshot of a network of long stringsand closed loops in the matter-dominated era.Figure taken from Ref. [17].

    of string intercommutations. The scaling regimeof string loops appears after a transient relaxationera, driven by a transient overproduction of stringloops with length close to the initial correlationlength of the string network. Subsequently, nu-merical [18] and analytical [19] studies supportedthe results of Ref. [17].

    Let me note that there are two approaches ofdeveloping numerical simulations of cosmic stringevolution. Either cosmic strings are modelled asidealised one-dimensional objects, or eld theo-retic calculations have been considered. In partic-ular, for the eld theoretic approach, the simplestexample of an underlying eld theory contain-ing local U(1) strings, namely the Abelian Higgsmodel, has been recently employed [20].

    5. String Thermodynamics

    It is well-known in string theory, that the de-generacy of string states increases exponentiallywith energy, namely

    d(E) eHE . (20)Hence, there is a maximum temperature Tmax =1/H, the Hagedorn temperature [21]. Let us con-

    sider, in the microcanonical ensemble, a systemof closed string loops in a three-dimensional box.Intersecting strings intercommute, but otherwisethey do not interact and are described by theGoto-Nambu equation of motion. The statisticalproperties of a system of strings in equilibrium arecharacterised by only one parameter, the energydensity of strings, , dened as = E/L3, with Lthe size of the cubical box. The behaviour of thesystem depends on whether it is at low or highenergy densities, and it undergoes a phase tran-sition at a critical energy density, the Hagedornenergy density H. Quantisation implies a lowercuto for the size of the string loops, determinedby the string tension . The lower cuto on theloop size is roughly 1/2, implying that the massof the smallest string loops is m0 1/2.

    For a system of strings at the low energy den-sity regime ( H), all strings are choppeddown to the loops of the smallest size, while largerloops are exponentially suppressed. Thus, forsmall enough energy densities, the string equilib-rium conguration is dominated by the masslessmodes in the quantum description. The energydistribution of loops, given by the number dn ofloops with energies between E and E + dE perunit volume, is [21,22]

    dn eEE5/2dE ( H) , (21)where = (5/2m0) ln(H/).

    However, as the energy density increases, moreand more oscillatory modes of strings get excited.In particular, once a critical energy density, H,is reached, long oscillatory string states begin toappear in the equilibrium state. The density atwhich this happens corresponds to the Hagedorntemperature. The Hagedorn energy density achieved when the separation between the small-est string loops is of the order of their sizes is approximately 2, and then the system under-goes a phase transition characterised by the ap-pearance of super-horizon (inntely long) strings.

    At the high energy density regime, the energydistribution of string loops is [21,22]

    dn = Am9/20 E

    5/2dE ( H) , (22)where A is a numerical coecient independentof m0 and of . Equation (22) implies that the

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 6890 75

  • mean-square radius R of the sub-horizon loops is

    R m3/20 E1/2 . (23)Hence the large string loops are random walksof step approximately m10 . Equations (22) and(23) imply

    dn = AR4dR ( H) , (24)where A is a numerical constant. Thus, at thehigh energy density regime, the distribution ofclosed string loops is scale invariant, since it doesnot depend on the cuto parameter. The totalenergy density in sub-horizon string loops is in-dependent of . Increasing the energy density ofthe system of strings, the extra energy E EH,where EH = HL

    3, goes into the formation ofsuper-horizon long strings, implying

    l = const ( H) , (25)where l denotes the energy density in super-horizon loops (often called in the literature as in-nite strings).

    Clearly, the above analysis describes the be-haviour of a system of strings of low or high en-ergy densities, while there is no analytic descrip-tion of the phase transition and of the interme-diate densities around the critical one, H.An experimental approach to the problem hasbeen proposed in Ref. [23] and later extended inRef. [24].

    The equilibrium properties of a system of cos-mic strings have been studied numerically inRef. [23]. The strings are moving in a three-dimensional at space and the initial string statesare chosen to be a loop gas consisting of the small-est two-point loops with randomly assigned po-sitions and velocities. This choice is made justbecause it oers an easily adjustable string en-ergy density. Clearly, the equilibrium state is in-dependent of the initial state. The simulationsrevealed a distinct change of behaviour at a crit-ical energy density H = 0.0172 0.002. For < H, there are no super-horizon strings, theirenergy density, l, vanishes. For > H, theenergy density in sub-horizon string loops is con-stant, equal to H, while the extra energy goes tothe super-horizon string loops with energy den-sity l = H. Thus, Eqs. (22) and (25) are

    valid for all > H, although they were de-rived only in the limit H. At the criticalenergy density, = H, the system of stringsis scale-invariant. At bigger energy densities, > H, the energy distribution of sub-horizonstring loops at dierent values of were found[23] to be identical within statistical errors, andwell-dened by a line dn/dE E5/2. Thus,for > H, the distribution of sub-horizon stringloops is still scale-invariant, but in addition thesystem includes super-horizon string loops, whichdo not exhibit a scale-invariant distribution. Thenumber distribution for super-horizon loops goesas dn/dE 1/E, which means that the totalnumber of super-horizon string loops is roughlylog(E EH). So, typically the number of longstrings grows very slowly with energy; for > Hthere are just a few super-horizon strings, whichtake up most of the energy of the system.

    The above numerical experiment has been ex-tended [24] for strings moving in a higher dimen-sional box. The Hagedorn energy density wasfound for strings moving in boxes of dimensional-ity dB = 3, 4, 5 [24]:

    H =

    0.172 0.002 for dB = 30.062 0.001 for dB = 40.031 0.001 for dB = 5


    Moreover, the size distribution of sub-horizonstring loops at the high energy density regime wasfound to be independent of the particular value of for a given dimensionality of the box dB. Thesize distribution of sub-horizon string loops wasfound [24] to be well dened by a line


    dE E(1+dB/2) , (27)

    where the space dimensionality dB was takenequal to 3, 4, or 5 . The statistical errors indi-cated a slope equal to (1 + dB/2) 0.2. Abovethe Hagedorn energy density the system is againcharacterised by a scale-invariant distribution ofsub-horizon string loops and a number of super-horizon string loops with a distribution which isnot scale invariant.

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 689076

  • 6. Cosmic String Gravity

    The gravitational properties of cosmic stringsare very dierent than those of non-relativisticlinear mass distributions. Straight cosmic stringsproduce no gravitational force in the surround-ing matter: 2 = 0, where stands for theNewtonian potential. Cosmic strings have rela-tivistic motion, implying that oscillating stringloops can be strong emitters of gravitational ra-diation. Since super-horizon cosmic strings havewiggles and small-scale structure due to string in-tercommutations, they are also sources of grav-itational radiation [25]. A gravitating string isdescribed by a coupled system of Einstein, Higgsand gauge eld equations, for which no exact solu-tion is known. We thus usually make two simpli-cations: we consider the cosmic string thicknessto be much smaller than any other relevant di-mension and the cosmic string gravitational eldto be suciently weak, so that linearised Einsteinequations can be used (for G 1).

    The geometry around a straight cosmic string islocally identical to that of at space-time, but thisgeometry is not globally Euclidean; the azimuthalcoordinate varies in the range [0, 2(1 4G)).Hence, the eect of a cosmic string is to intro-duce an azimuthal decit angle , whose mag-nitude is determined by the symmetry breakingscale Tc leading to the cosmic string formation,namely = 8G. Thus, the string metricds2 = dt2dz2dr2(18G)r2d2 describesa conical space leading to interesting observa-tional eects on the propagation of light (i.e.,double images of light sources located behind cos-mic strings) and of particles (i.e., discontinuousDoppler shift eects). The centre-of-mass veloc-ity u of two particles, moving towards a string atthe same velocity v reads

    u = v sin(/2)[1 v2 cos2(/2)]1/2 . (28)Considering that one of the particles carries alight source, while the other one is an observer,one realises that the observer will detect a discon-tinuous change in the frequency of light, givenby [26]


    v1 v2 . (29)

    This discontinuous change in the frequency has itsorigin in the Doppler shift: particles start mov-ing towards each other, decreasing their distance,once the line connecting the particles crosses thestring.

    In the framework of gravitational instability,topological defects in general and cosmic stringsin particular, oered an alternative to the ina-tionary paradigm for the origin of the initial uc-tuations leading to the observed large-scale struc-ture and the measured anisotropies of the CMBtemperature anisotropies. The angular powerspectrum of CMB is expressed in terms of thedimensionless coecients C, in the expansion ofthe angular correlation function in terms of theLegendre polynomials P reads


    TT (n)T



    (nn=cos )



    (2 + 1)CP(cos)W2 , (30)

    where W stands for the -dependent windowfunction of the particular experiment. Equation(30) compares points in the sky separated by anangle . The value of C is determined by uctu-ations on angular scales of the order of /. Theangular power spectrum of anisotropies observedtoday is usually given by the power per logarith-mic interval in , plotting ( + 1)C versus .

    To nd the power spectrum induced by topo-logical defects, one has to solve in Fourier space,for each given wave vector k, a system of linearperturbation equations with random sources:

    DX = S , (31)where D denotes a time dependent linear dif-ferential operator, X is a vector which containsthe various matter perturbation variables, and Sis the random source term, consisting of linearcombinations of the energy momentum tensor ofthe defect. For given initial conditions, Eq. (31)can be solved by means of a Greens function,G(, ). To compute power spectra or, more gen-erally, quadratic expectation values of the form

    6Equation (30) holds only if the initial state for cosmo-logical perturbations of quantum-mechanical origin is thevacuum [27].

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 6890 77

  • Xj(0,k)Xm(0,k), one has to calculate

    Xj(0,k)Xl (0,k) = 0





    d Gln( ,k)Sm(,k)Sn( ,k) .(32)

    To compute power spectra, one shouldknow the unequal time two-point correlatorsSm(,k)Sn( ,k) in Fourier space, calculatedby means of heavy numerical simulations.

    The rst determinations of the CMB powerspectrum from cosmic strings were based on theassumption that cosmic strings are of innites-imal width. They were thus realised by eitheremploying Nambu-Goto simulations of connectedstring segments, or a model involving a stochas-tic ensemble of unconnected segments. However,a number of questions there have been later raisedregarding the accuracy of these approaches. Morerecently, the CMB power spectrum contributionfrom cosmic strings has been addressed [28] usingeld-theoretic simulations of the Abelian Higgsmodel, the simplest example of an underlyingeld theory with local U(1) strings. All ap-proaches agree on the basic form of the cosmicstring power spectrum, namely a power spectrumwith a roughly constant slope at low multipoles,rising up to a single peak, and consequently de-caying at small scales. This result is common ofall models, in which uctuations are generatedcontinuously and evolve according to inhomoge-neous linear perturbation equations.

    In topological defects models, uctuations areconstantly generated by the non-linear defect evo-lution. This characteristic, combined with thefact that the random initial conditions of thesource term of a given scale leak into other scales,destroy perfect coherence. The incoherent aspectof active perturbations aects the structure ofsecondary oscillations, namely secondary oscilla-tions may get washed out. Thus, in topologicaldefects models, incoherent uctuations lead to asingle bump at smaller angular scales (larger ),than those predicted within any inationary sce-nario.

    The cosmic string CMB power spectrum wasfound [28] to have a broad peak at 150 400.Decomposing the power spectrum into scalar,

    vector and tensor modes, it was shown [28] thatthe origin of this broad peak lies in both the vec-tor and scalar modes, which peak at 180and 400, respectively. This analysis con-cluded [28] that the cosmic string power spec-trum is dominated by vector modes for all butthe smallest scales.

    The position and amplitude of the acousticpeaks, as found by the CMB measurements, are indisagreement with the predictions of topologicaldefect models. As a consequence, CMB measure-ments rule out pure topological defect models ingeneral, and cosmic strings in particular, as theorigin of initial density perturbations leading tothe observed structure formation.

    Since cosmic strings are expected to be gener-ically formed in the context of SUSY GUTs, oneshould consider mixed perturbation models wherethe dominant role is played by the inaton eldbut cosmic strings have also a contribution, smallbut not negligible. Restricting ourselves to theangular power spectrum, we can remain in thelinear regime. In this case,

    C = CI

    + (1 )CS , (33)where CI and C


    denote the (COBE normalised)Legendre coecients due to adiabatic inatonuctuations and those stemming from the cosmicstring network, respectively. The coecient inEq. (33) is a free parameter giving the relativeamplitude for the two contributions. Comparingthe C, calculated using Eq. (33) where C


    istaken from a generic inationary model and CSfrom numerical simulations of cosmic string net-works with data obtained from the most re-cent CMB measurements, one gets that a cosmicstring contribution to the primordial uctuationshigher than 14% is excluded up to 95% condencelevel [29].

    Let us now return to F- and D-term hybridination and investigate the constraints on thefree parameters of the model (namely massesand couplings) so that the cosmic string con-tribution to the CMB data is within the al-lowed limits imposed from recent CMB measure-ments. Considering only large angular scales onecan get the contributions to the CMB temper-ature anisotropies analytically. The quadrupole

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 689078

  • anisotropy has one contribution coming from theinaton eld, and one contribution coming fromthe cosmic string network. Fixing the number ofe-foldings to 60, the inaton and cosmic stringcontribution to the CMB depend on the parame-ters of the model. For F-term ination the cosmicstring contribution to the CMB data is consistentwith CMB measurements provided [4]


  • Apart the temperature power spectrum, impor-tant constraints on cosmic string scenarios mightalso arise in the future from measurements of thepolarisation of the CMB photons. More precisely,the B-polarisation spectrum oers an interestingwindow on cosmic strings [30] since ination hasonly a weak contribution. Scalar modes maycontribute to the B-mode only via the gravita-tional lensing of the E-mode signal, with a secondinationary contribution coming from the sub-dominant tensor modes.

    Cosmic strings can also become apparentthrough their contribution in the small-angleCMB temperature anisotropies. More precisely,at high multipoles (small angular resolution),the mean angular power spectrum of string-induced CMB temperature anisotropies can bedescribed [31] by , with 0.889. Thus,a non-vanishing cosmic string contribution to theoverall CMB temperature anisotropies may dom-inate at high multipoles (small angular scales).In an arc-minute resolution experiment, stringsmay be observable [31] for G down to 2 107.

    Cosmic strings should also induce deviationsfrom Gaussianity. On large angular scales suchdeviations are washed out due to the low stringcontribution, however on small angular scales, op-timal non-Gaussian string-devoted statistical es-timators may impose severe constraints on a pos-sible cosmic string contribution to the CMB tem-perature anisotropies.

    Finally, let me emphasise that one should keepin mind that all string-induced CMB temperatureanisotropies were performed for Abelian strings inthe zero thickness limit with reconnection prob-ability equal to unity and winding number equalto one. Even though in any model where uctu-ations are constantly induced by sources (seeds)having a non-linear evolution, the perfect coher-ence which characterises the inationary inducedspectrum of perturbations gets destroyed [32],there is still no reason to expect that quantita-tively the results found for conventional cosmicstring models will hold in more general cases.

    7. Superconducting Cosmic Strings

    Before nishing this brief review on cosmicstrings, let me mention the case of superconduct-ing strings [33], in the sense that in a large class ofhigh energy physics theories, strings have similarelectromagnetic properties as those of supercon-ducting wires. Such objects carry large electriccurrents and hence their interaction with the cos-mic plasma can lead to a variety of distinct as-trophysical eects.

    Cosmic strings are characterised as supercon-ductors if electromagnetic gauge invariance is bro-ken inside the strings, a situation which can oc-cur for instance when a charged scalar eld de-velops a non-zero expectation value in the vicin-ity of the string core 7. Superconducting stringsappear also in models with fermions, which ac-quire masses through a Yukawa coupling to theHiggs eld of the strings. Thus, depending onthe considered model we can have bosonic orfermionic string superconductivity. In the rstcase, bosons can condensate and acquire a non-vanishing phase gradient, while in the second one,fermions may propagate in the form of zero modesalong the string.

    Applying an electric eld on a superconductingstring, the string will develop a growing electriccurrent according to


    dt ce


    E , (35)

    where E stands for the eld component along thestring and e denotes the elementary charge.

    In the case of fermionic superconductivity,fermions are massless inside the string, whereasthey have a nite mass, m, outside the string 8.Under the eect of an electric eld, a current Jresults, growing in time until it reaches a criticalvalue Jc emc2/, when the particles inside thestring, moving at relativistic speeds, have su-cient energy to leave the string. Thus, when thestring current reaches its critical value, Jc, parti-

    7Note that also a vector eld, whose ux is trapped insidea non-Abelian string can lead to a superconducting string.8The fermion mass is model-dependent, but it is boundedfrom above by the symmetry breaking scale of the string.

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 689080

  • cles get produced at a rate

    n eE/ , (36)where n stands for the number of fermions perunit length. Note that Jc, even though model-dependent, it does not exceed a maximum value,given by Jmax e(c3/)1/2. Depending on theirenergy scale, superconducting strings may carryhuge currents. In the case of bosonic supercon-ductivity, the model-dependent critical current Jc determined by the energy scale at which scaleinvariance is broken is again bounded by Jmax,dened as above.

    Superconducting strings can also develop grow-ing currents in magnetic elds, according to


    dt e


    vB , (37)

    where v is the speed of the moving string segmentin a magnetic eld B. For a string loop carryingsuciently large currents, electromagnetic radia-tion can overtake gravitational radiation, becom-ing the dominant energy loss mechanism.

    Superconducting string loops may be problem-atic in cosmology. For a current-carrying loop,the energy per unit length is not equal to thetension, their dierence equals the string cur-rent. Such a loop can rotate and the resultingcentrifugal force which can be expected to bevery much stronger than the inecient magneticspring repulsion eect may balance the ten-sion. When a rotating string loop reaches an equi-librium state dened by the balance betweenthe string tension and the centrifugal force it iscalled a vorton [34]. Vortons can be formed at, orsoon after, the phase transition followed by SSBleading to the string formation, and they possessa net charge as well as a current. If vortons arestable (certainly a model-dependent issue [35]),they will scale as matter in the Universe, domi-nating over its energy density. In this sense, or-tons may constrain models for superconductingstrings.

    8. Cosmic Superstrings

    The recent interplay between superstring the-ory and cosmology has led to the notion of cosmic

    superstrings [3], providing the missing link be-tween superstrings and their classical analogues.

    The possible astrophysical role of superstringshas been advocated already more than twentyyears ago. More precisely, it has been pro-posed [36], that superstrings of the O(32) andE8E8 string theories are likely to generatestring-like stable vortex lines and ux tubes.However, in the context of perturbative stringtheory, the high tension (close to the Planck scale)of fundamental strings ruled them out [36] as po-tential cosmic string candidates. Luckily, thispicture has changed in the framework of brane-world cosmology, which oers an elegant realisa-tion of nature within string theory. Within thebrane-world picture, all standard model particlesare open string modes. Each end of an openstring lies on a brane, implying that all standardmodel particles are stuck on a stack of Dp-branes,while the remaining p 3 of the dimensions arewrapping some cycles in the bulk. Closed stringmodes (e.g., dilaton, graviton) live in the high-dimensional bulk. Brane interactions lead to un-winding and thus evaporation of higher dimen-sional Dp-branes. We are eventually left withD3-branes one of which could indeed play therole of our Universe [37] embedded in a (9+1)-dimensional bulk and cosmic superstrings (one-dimensional D-branes, called D-strings, and Fun-damental strings, called F-strings).

    Brane annihilations provide a natural mecha-nism for ending ination. To illustrate the forma-tion of cosmic superstrings at the end of brane in-ation, let us consider a Dp-Dp brane-anti-branepair annihilation to form a D(p 2) brane. Eachparent brane has a U(1) gauge symmetry andthe gauge group of the pair is U(1)U(1). Thedaughter brane possesses a U(1) gauge group,which is a linear combination, U(1), of theoriginal two U(1)s. The branes move towardseach other and as their inter-brane separation de-creases below a critical value, the tachyon eld,which is an open string mode stretched betweenthe two branes, develops an instability. Thetachyon couples to the combination U(1). Therolling of the tachyon eld leads to the decay ofthe parent branes. Tachyon rolling leads to spon-taneously symmetry breaking, which supports de-

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 6890 81

  • fects with even co-dimension. So, brane annihi-lation leads to vortices, D-strings; they are cos-mologically produced via the Kibble mechanism.The other linear combination, U(1)+, disappearssince only one brane remains after the brane colli-sion. The U(1)+ combination is thought to disap-pear by having its uxes conned by fundamen-tal closed strings. Such strings are of cosmolog-ical size and they could play the role of cosmicstrings [38]; they are referred to in the literatureas cosmic superstrings [39].

    9. Dierences between Cosmic Strings and

    Cosmic Superstrings

    Cosmic superstrings [3], even though cosmolog-ically extended, are quantum objects, in contrastto solitonic cosmic strings which are classical ob-jects. Hence, one expects a number of dierencesto arise as regarding the properties of the twoclasses of objects. As I have earlier discussed,the probability that a pair of cosmic strings willreconnect, after having intersect, equals unity.The reconnection probability for cosmic super-strings is however smaller (often much smaller)than unity. The corresponding intercommutationprobabilities are calculated in string perturbationtheory. The result depends on the type of stringsand on the details of compactication. For fun-damental strings, reconnection is a quantum pro-cess and takes place with a probability of order g2s(where gs denotes the string tension). It can thusbe much less than one, leading to an increaseddensity of strings [40], implying an enhancementof various observational signatures. The recon-nection probability is a function of the relativeangle and velocity during the collision. One maythink that strings can miss each other, as a resultof their motion in the compact space. Dependingon the supersymmetric compactication, stringscan wander over the compact dimensions, thusmissing each other, eectively decreasing their re-connection probability. However, in realistic com-pactication schemes, strings are always connedby a potential in the compact dimensions [41].The value of gs and the scale of the conning po-tential will determine the reconnection probabil-ity. Even though these are not known, for a large

    number of models it was found [41] that the recon-nection probability for F-F collisions lies in therange between 103 and 1. The case of D-D col-lisions is more complicated; for the same modelsthe reconnection probability is anything between0.1 to 1. Finally, the reconnection probability forF-D collisions can vary from 0 to 1.

    Brane collisions lead not only to the formationof F- and D-strings, they also produce boundstates, (p, q)-strings, which are composites of pF-strings and q D-strings [42]. The presence ofstable bound states implies the existence of junc-tions, where two dierent types of string meetat a point and form a bound state leading awayfrom that point. Thus, when cosmic superstringsof dierent types collide, they can not intercom-mute, instead they exchange partners and forma junction at which three string segments meet.This is just a consequence of charge conserva-tion at the junction of colliding (p, q)-strings. Forp = np and q = nq, the (p, q) string is neutrallystable to splitting into n bound (p, q) strings.The angles at which strings pointing into a ver-tex meet, is xed by the requirement that therebe no force on the vertex. In general, a (p, q) anda (p, q) string will form a trilinear vertex witha (p + p, q + q) or a (p p, q q) string. Thisleads certainly to the crucial question of whethera cosmic superstring network will reach scaling,or whether it freezes leading to predictions incon-sistent with our observed Universe universe.

    The tension of solitonic strings is set from theenergy scale of the phase transition, followed bya spontaneously broken symmetry, which left be-hind these defects as false vacuum remnants. Cos-mic superstrings however span a whole range oftensions, set from the particular brane inationmodel. The tension of F-strings in 10 dimensionsis F = 1/(2

    ), and the tension of D-strings isD = 1/(2

    gs), where gs stands for the stringcoupling. In 10 at dimensions, supersymmetrydictates that the tension of the (p, q) bound statesreads

    (p,q) = F

    p2 + q2/g2s . (38)

    Individually, the F- and D-strings are 12 -BPS (Bogomolnyi-Prasad-Sommereld) objects,which however break a dierent half of the su-

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 689082

  • persymmetry each. Equation (38) represents theBPS bound for an object carrying the chargesof p F-strings and q D-strings. Note that theBPS bound is saturated by the F-strings, (p, q) =(1, 0), and the D-strings, (p, q) = (0, 1). Let mealso make the remark that the string tension forstrings at the bottom of a throat is dierent fromthe (simple) expression given in Eq. (38) and itdepends on the choice of ux compactication.

    Consider an F- and a D-string, both lying alongthe same axis. The total tension of this cong-uration is (g1s + 1)/(2

    ), which exceeds theBPS bound; thus the conguration is not super-symmetric. It can however lower its energy, if theF-string breaks, its ends being attached to the D-string. Since the end points can then move o atinnity, leaving only the D-string behind, a uxwill run between the end points of the F-string.Now the tension reads (g1s +O(gs))/(2), andhence the nal state represents a D-string with aux, which is a supersymmetric state.

    10. Cosmic Superstring Evolution

    The evolution of cosmic superstring networks,is a complicated issue, which has been addressedby numerical [40,4345], as well as analytical [46]approaches.

    The rst numerical attempt [40], studying inde-pendent stochastic networks of D- and F-stringsin a at space-time, has shown that the charac-teristic length scale , giving the typical distancebetween the nearest string segments and the typ-ical curvature of strings, grows linearly with time

    (t) t ; (39)the slope depends on the reconnection proba-bility P , and on the energy of the smallest al-lowed loops (i.e., the energy cuto). For recon-nection (or intercommuting) probability in therange 103

  • where tan = q/(pgs). The balance conditionsfor three strings imply that when an F-stringends on a D-string, it causes it to bend at anangle set by the string coupling; on the otherside of the junction there is a (1, 1) string. Con-sider a junction of three strings, with coordi-nates x(, t) , tension and parameter lengthsL1(t), L2(t), L3(t),which are joined at a junctionand whose other end terminate on parallel branes.The action for this conguration reads 10 [46]

    S = 3







    dtl (x(t, L(t))) xjunc(t)) , (43)

    where the rst part stands for the Nambu-Gototerms for the three strings, and l denote the La-grange multipliers to describe the junction, lo-cated at position xjunc. From Eq. (43) one canderive the equations of motion as well as the en-ergy conservation. One can easily check that

    1(1 L1)1 + 2 + 3


    M1(1 c23)M1(1 c23) + M2(1 c13) + M3(1 c12) ,

    1L1 + 2L2 + 3L3 = 0 , (44)

    and cyclic permutations. Note that M1 =21 (2 3)2 (and cyclic permutations givingM2,M3) and cij = a

    i(t Li(t)) aj(t Lj(t)).

    Equation (44) implies that the rate of creation ofnew string must balance the disappearance of oldone. Thus, for an F-string with 1 = 1, and aD-string with 2 = 1/gs, the FD-bound state hastension 3 =

    1 + 1/g2s = 1/gs + gs/2 + O(g2s ).

    Since the angle goes to /2 in the limit of zerostring coupling, we conclude that in the small gs-limit, the length of the F-string remains constant,while the length of the D-string decreases and thelength of the FD-bound state increases. This re-sult has been recently conrmed from numericalexperiments [45].

    10Note that in principle, cosmic superstring dynamicsought to be studied using the Dirac-Born-Infeld action,the low-energy eective action for many varieties of stringsarising in the context of string theory.

    To shed some light on the evolution of cosmicsuperstring networks, a number of numerical ex-periments have been conducted, each of them ata dierent level of approximation. One shouldkeep in mind that the initial conguration de-pends on the particular brane ination scenario,while a realistic network should contain stringswith junctions and allow for a spectrum of possi-ble tensions.

    I will briey describe the approach and ndingsof one of these numerical approaches [44], whichI consider more realistic than others. The aimof that study was to build a simple eld theorymodel of (p, q) bound states, in analogy with theAbelian Higgs model used to investigate the prop-erties of solitonic cosmic string networks, and tostudy the overall characteristics of the networkusing lattice simulations. Two models were in-vestigated, one in which both species of stringhave only short-range interactions and anotherone in which one species of string features long-range interactions. We thus modelled the networkwith no long-range interactions using two sets ofelds, complex scalars coupled to gauge elds,with a potential chosen such that the two types ofstrings will form bound states (see, Fig. 3). In thisway junctions of 3 strings with dierent tensionwere successfully modelled. In order to introducelong-range interactions we considered a networkin which one of the scalars forms global strings.This is important if the strings are of a non-BPSspecies. For example, for cosmic superstrings atthe bottom of a Klebanov-Strassler throat the F-string is not BPS while the D-string is.

    More precisely, the (p, q) string network wasmodelled [44] using two sets of Abelian Higgselds, , . In the case that both species of cos-mic strings are BPS, the model is described bythe action [44]:

    S =


    [ 1

    4F 2 1

    2(D) (D



    ( 21


    14H2 1

    2(D) (D



    ( 22

    )2 ], (45)

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 689084

  • where the covariant derivative D is dened by

    D = ie1A ,D = ie2C . (46)For clarity, we label the eld as Higgs andthe eld as axion, even though both eldsare Higgs-like. The scalars are coupled to theU(1) gauge elds A and C, with coupling con-stants e1 and e2 and eld strength tensors F =A A and H = C C, respec-tively. The scalar potentials are parametrised bythe positive constants 1, 1 and 2, 2, respec-tively. In the case that one species of string isnon-BPS, we remove the second gauge eld bysetting e2 = 0. In this way, this species of string isrepresented by the topological defect of a complexscalar eld with a global U(1) symmetry. Notethat such defects are characterised by the exis-tence of long-range interactions as opposed tolocal strings in which all energy density is con-ned within the string, so that local strings haveonly gravitational interactions implying dier-ent consequences for the evolution of the network.

    Figure 3. Left: Bound states for local-local (p, q)strings. Right: Bound states for local-global (p, q)strings. Figure taken from Ref. [44].

    Thus, dierent components of the (p, q) stateare expected to exhibit dierent types of long-range interactions. The evolution of the stringnetworks suggested that the long-range interac-tions have a much more important role in thenetwork evolution than the formation of bound

    states. In the local-global networks the boundstates tend to split as a result of the long-rangeinteractions, resulting in two networks that evolvealmost independently. The formation of short-lived bound states and their subsequent splittingonly increases the small-scale wiggliness of the lo-cal strings. In the case of a local-local network,the absence of long-range interactions allows thebound states to be much longer-lived and signif-icantly inuences the evolution of the string net-work [44]. The most convincing evidence comesfrom analysing the reverse problem [44], namelythat of a bound state splitting as a result ofthe long-range interactions between strings, pre-sented in Fig. 4. Only in the absence of long-rangeinteractions the strings remain in the (1, 1) statethroughout their entire evolution.





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    Figure 4. The total physical volume of the simu-lation box occupied by Higgs strings (green), ax-ion strings (red), and their bound states (blue).The left panels refer to local-global networks,while the right ones to local-local networks start-ing from the same initial conditions as the local-global ones. Figure taken from Ref. [44].

    Let us now investigate more thoroughly the is-sue of scaling. The evolution of F-, D-strings and

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 6890 85

  • their bound states is a rather complicated prob-lem, which necessitates both numerical as well asanalytical investigations. As I have already men-tioned, junctions may prevent the network fromachieving a scaling solution, invalidating the cos-mological model leading to their formation. Fol-lowing the approach of Ref. [44], numerical simu-lations [45], achieving control over the initial pop-ulation of bound states, found clear evidence forscaling of all three components p F-strings, qD-strings and their (p, q) bound states of thenetwork, independently of the chosen initial con-gurations, while they concluded that the exis-tence of bound states eects the evolution of thenetwork. In Fig. 5 we show the string correla-tion length for the Higgs and axion elds, as wellas for their bound states, as a function of time.The initial conguration is a local-global networkwith a large amount of bound states. The corre-









    gth 26





    30 40 Time 60 70 10













    30 40 Time 60 70












    30 40 Time 60 70

    Figure 5. The Higgs (left), axion (middle), andbound state (right) string correlation length as afunction of time. The network is a local-globalone. The data and linear ts for the two regimesare shown. Figure taken from Ref. [45].

    sponding plots for local-local networks are drawnin Fig. 6. Clearly, there is convincing evidence forscaling of the three components of the network for











    30 40 Time 60 70 24










    30 40 Time 60 70









    gth 40





    30 40 Time 60 70

    Figure 6. The same as Fig. 5, but for a local-localnetwork. Figure taken from Ref. [45].

    both networks. This scaling is characterised witha distinct change of the correlation length slopeduring the network evolution. Note that the re-sult holds even in the case of networks with smallamounts of bound states.

    Moreover, these numerical experiments haveshown that for (p, q) strings there is a supplemen-tary energy loss mechanism, in addition to thechopping o of loops; it is this new mechanismthat allows the network to scale. More precisely,the additional energy loss mechanism is the for-mation of bound states, whose length increases,lowering the overall energy of the network.

    11. Cosmic Superstrings: A window into

    String Theory

    Cosmic superstrings have gained a lot of in-terest, the main reason being that they can of-fer a large (and possibly unique) window intostring theory, and in particular shed some lighton the appropriate (if any) stringy description ofthe Universe. Since they interact with StandardModel particles only via gravity, their detectioninvolves their gravitational interactions. Cosmicsuperstrings, in an analogy to their solitonic ana-logues, can lead to a variety of astrophysical sig-natures, like gravitational waves, ultra high en-

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 689086

  • ergy cosmic rays, and gamma ray bursts.At this point, let me however emphasise that

    given the complexity of the dynamics of a cosmicstring network, which we certainly do not fullyunderstand, and the model-dependent initial con-guration, any theoretical estimations of the ob-servational signatures of cosmic superstrings haveto be taken with caution. Note that even the su-perstring tension depends on the considered ina-tionary scenario within a particular brane-worldcosmological model.

    Gravitational waves is one of the main exploredavenues [47], in which case three channels of emis-sion have been identied. Radiation can be emit-ted by cusps, kinks, and/or from the reconnec-tion process itself. Cusps, where momentarilythe string moves relativistically, have played acrucial role in discussing the radiation emittedfrom (ordinary) cosmic strings 11. Kinks, result-ing from cosmic string collisions and subsequentreconnection, are basically replaced in the caseof cosmic superstrings by junctions. Finally, theradiation emitted from the reconnection processitself, which a sub-dominant process in the case ofcosmic strings, may not be negligible in the caseof cosmic superstrings because of the small re-connection probability P . In the scaling regime,the density of long strings goes like 1/P [40], im-plying that the number of reconnection attemptsgoes like 1/P2, and hence the number of success-ful reconnections is approximately 1/P . Very re-cently, the gravitational waveform produced bycosmic superstring reconnections has been calcu-lated [49]. Comparing the obtained result to thedetection threshold for current and future grav-itational wave detectors, it was concluded [49]that neither bursts nor the stochastic gravita-tional background, produced during the cosmicsuperstring reconnection process, would be de-tectable by Advanced LIGO. Thus, the most rele-vant process for gravitational waves emitted fromcosmic superstrings turns out to be through theircusps. Hence, one should estimate the abundancyof cusps in cosmic superstrings with junctions.

    11Even though one has to keep in mind that the number ofcusps in a realistic cosmi string network has not been esti-mated, while preliminary numerical studies indicate thatit may be rather low [48].

    Following simple geometric arguments, it hasbeen recently shown [50] that strings ending on D-branes can indeed lead to cusps, in an analogousway as cusps in ordinary cosmic strings. In par-ticular, cusps would be a generic feature of an F-string ending on two (parallel and stationary) D-strings. Hence, pairs of FD-string junctions, suchas those that they would form after intercommu-tations of F- and D-strings, generically containcusps. This result opens up a new energy lossmechanism for the network, in addition to the for-mation and subsequent decay of closed loops andthe formation of bound states [45]. Phenomeno-logical consequences of cusps from junctions oncosmic superstrings will be most signicant atearly times, namely towards the end of brane in-ation, since then the typical separation of heavystrings is small as compared to the length of F-strings stretched between them [50].

    12. Cosmic Superstring Thermodynamics

    One has to extend previous studies of stringthermodynamics in the case of cosmic superstringnetworks, characterised by the existence of (p, q)bound states and dierent string tensions. Re-cently, the Hagedorn transition of strings withjunctions has been investigated [51], in the con-text of a simple model with three dierent typesand tensions of string, following an eective eldtheory approach. More precisely, the authors ofRef. [51] translated the thermodynamics of stringnetworks with junctions into the thermodynamicsof a set of interacting dual elds. Thus, the Hage-dorn transition of the strings becomes a transitionof the elds.

    In this approach, the equilibrium statisticalmechanics of cosmic superstring networks havebeen studied [51], by extending known meth-ods for describing quark deconnement. It wasfound [51] that as the system is heated, thelightest strings are the rst ones to undergo aHagedorn transition; the existence of junctionsdoes not aect the occurrence of the transition.The system is also characterised by a second,higher, critical temperature above which longstring modes of all tensions and junctions, do ex-ist. The existence of multiple tensions indicates

    M. Sakellariadou / Nuclear Physics B (Proc. Suppl.) 192193 (2009) 6890 87

  • the appearance of multiple Hagedorn transitions.

    13. Conclusions

    In these lectures, I have summarised our cur-rent understanding on the physics of cosmicstrings and cosmic superstrings. I have discussedtheir formation, evolution, statistical mechan-ics and astrophysical/cosmological consequences.This is a topic of active research at present, re-lating fundamental theoretical ideas with experi-mental and observational facts.

    On the one hand, any successful cosmologicalscenario, such as the inationary paradigm, mustbe inspired from a fundamental theory. On theother hand, any successful high energy physicstheory, such as string theory or supersymmet-ric grand unied theories, must be tested againstdata; the only available laboratory for the re-quired energy scales, is indeed the early Universe.

    Ination within brane-world cosmological mod-els leads naturally to cosmic superstrings. Ina-tion within supersymmetric grand unied theo-ries leads generically to cosmic strings, the soli-tonic analogues of cosmic superstrings. The studyof these objects is interesting by itself. In addi-tion, cosmic (super)strings may provide an ex-planation for the origin of a variety of astrophysi-cal/cosmological observations; they may also oera test (often a unique one) of fundamental theo-ries of physics, thus shedding some light about theappropriate stringy description of the Universe.


    It is a pleasure to thank the organisers of theESF Summer School in High Energy Physics andAstrophysics Theory and Particle Physics: theLHC perspective and beyond, which took placeat the Carge`se Institute of Scientic Studies inCorsica, for inviting me to present these lecturesin such a beautiful and stimulating environment.


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