Revisiting light stringy states in view of the 750 GeV diphoton excess

1.1 MB·

by  Pascal Anastasopoulos

TUW-16-01
PreprinttypesetinJHEPstyle-HYPERVERSION
ROM2F/2016/01
Revisiting light stringy states
in view of the 750 GeV diphoton excess
6
1
0
2 Pascal Anastasopoulos1∗, Massimo Bianchi2†,

p
e 1 Technische Universit¨at Wien
S
Institut fu¨r Theoretische Physik

8
A-1040 Vienna, Austria

] 2 Dipartimento di Fisica and Sezione I.N.F.N.
h
t Universit`a di Roma “Tor Vergata”
-
p Via della Ricerca Scientifica 1, 00133 Roma, Italy
e
h
[
Abstract: We investigate light massive string states that appear at brane inter-
2
v sections. They replicate the massless spectrum in a richer fashion and may be para-
4
metrically lighter than standard Regge excitations. We identify the first few physical
8
5 states and determine their BRST invariant vertex operators. In the supersymmetric
7
0 case we reconstruct the super-multiplet structure. We then compute some simple
.
1 interactions, such as the decay rate of a massive scalar or vector into two massless
0
fermions. Finally we suggest an alternative interpretation of the 750 GeV diphoton
6
1 excess at LHC in terms of a light massive string state, a replica of the Standard
:
v Model Higgs.
i
X
r
a
∗[email protected]
†[email protected]
Contents
1. Introduction 2
2. Setup 4
3. Character valued partition function 5
4. States and vertex operators 7
4.1 Gauge sector 7
4.2 Massless sector 8
4.3 Massive multiplets 9
4.3.1 Massive multiplets with α(cid:48)m2 = a 10
1,2
4.3.2 Massive multiplets with α(cid:48)m2 = a 11
3
5. SUSY transformations 14
5.1 Massless multiplets 14
5.2 Massive multiplets 14
6. Interactions 17
7. Conclusions and outlook 22
A. Theta functions and Riemann identity 23
B. Useful OPE’s 23
C. State-Vertex operator dictionary 25
– 1 –
1. Introduction
Starting from the pioneering works on the systematics of open string theories [1–10],
that led to the first chiral model in D = 4 [11], intersecting D-brane models [12,13]
proved to be a promising class of models that allow to embed the Standard Model
or extensions thereof in String Theory1.
One of the most exciting properties of this class of models is that they allow for
a very low string scale, even of order of a few TeV’s [19–21], leading to interesting
signatures detectable at particle colliders such as the LHC or in cosmological ob-
servations such as Planck. Particular attention has been devoted to anomalous Z(cid:48)
physics (see, e.g. [22–35]), Kaluza Klein states (see, e.g. [36–42]), or purely stringy
signatures (see, e.g. [43–56])2.
A series of papers [46,48–50] focused on tree-level string scattering amplitudes
containing massless bosons and fermions that can be identified with the SM fields.
Amplitudes containing at most two chiral fermions were shown to exhibit a universal
behavior independently of the specifics of the compactification. The poles of these
scattering amplitudes correspond to the exchange of Regge excitations of the SM
gauge bosons, whose mass scales like the string mass. Due to the universal behavior
of these amplitudes, which in turn gives them a predictive power, LHC is able to
constrain the string scale to be above 4.5 TeV [59–61].
M
~ ~~
θ ψ, ψ, ψ ...
N
Figure 1: Towers of states whose masses are multiples of the intersection angle live at
D-brane intersections.
Inthispaper, ourattentiongoestowardsadifferentdirection. Attheintersection
of two D-brane stacks there exists a whole tower of stringy excitations with the same
quantum numbers [62–65]. The lowest state can be massless (for a specific value
of the intersecting angles) and the rest are massive with masses proportional to the
1For recent reviews on D-brane model building as well as specific MSSM D-brane constructions,
see [14–18] and references therein.
2For recent reviews, see [57,58].
– 2 –
˜ ˜˜
angle of the intersection in which they live (see fig. 1). Therefore if we call ψ, ψ, ψ
etc the states living at the intersection, we have
M2 = 0 , α(cid:48)M2 = θ/π , α(cid:48)M2 = 2θ/π ... (1.1)
ψ ψ˜ ψ˜˜
where θ denotes the intersection angle between these two D-brane stacks. For small
intersection angle those light stringy states can be significantly lighter than the first
Regge excitations of the gauge bosons and are expected to be observed prior to the
latter. In figure 2 we depict a D-brane SM realization3. While the SM gauge bosons
live on the world volume of the D-brane stacks the SM fermions are localized at the
intersections of different D-brane stacks. At each intersection there exist a tower of
massive stringy excitations that can be significantly lighter than the string scale for
some regions in the parameter space.
c
g
b
a
~ ~~ ~ ~~
u, u, u...
Q, Q, Q...
~ ~~
d, d, d...
U(1)
Z,W+_ c
~ ~~ U(1)d ~ ~~
l, l, l... e, e, e...
d
Figure 2: Local D-brane realization of the Standard Model. Each field is accompanied by
a whole tower of massive excitations.
Such towers of states are different from standard string resonances or Kaluza-
Klein excitations, for two main reasons.
• First of all, for standard Regge resonances the mass square grows linearly with
the level and is the same for each ground-state. However, in this D-brane setup
one has for each massless fermion a separate tower of stringy excitations with a
different mass spacing. That gives a larger variety for the masses of the excited
SM fields.
3FororiginalworkonlocalD-braneconfigurations,see[66,67]. Forasystematicanalysisoflocal
D-brane configurations, see [68–71].
– 3 –
• Second main difference is that Kaluza-Klein scenarios forbid decays which vi-
olate momentum conservation in the compact dimension. These decays, even
though with a small rate (depending on several moduli), are allowed in inter-
secting D-brane setups giving a lot of space for interesting phenomenology.
Aim of the present investigation is to carefully analyse the spectrum of lowest ly-
ing massive states in the case of supersymmetric intersections and provide the vertex
operators for the first few excitations. We will also discuss some simple interactions
and offer an alternative interpretation of the putative LHC signal at 750 GeV as a
light massive open string state living at a D-brane intersection, a replica of the SM
Higgs. In future works we are planning to use the toolkit developed here in order
to study the peculiar phenomenology of these states and give further support to our
present preliminary comments on the diphoton excess.
The paper is organized as follows. After setting up the stage in Section 2, we
computethecharacter-valuedpartitionfunctioninSection3andidentifythephysical
states together with their helicity. Then, in Section 4, we determine the relevant
vertex operators and impose the conditions for BRST invariance and in Section 5
we reconstruct the super-multiplet structure. Finally, in Section 6 we compute some
simple interactions, such as the decay rate of a massive scalar or vector into two
masslessfermions, and inSection7wepresent ourconclusionsandoutlook, including
comments on our alternative interpretation of the 750 GeV diphoton excess at LHC
in terms of a massive but light Higgs replica, that couples to two photons thanks to
massive string exchange. Several appendices contain useful formulae for the twist-
fields, their OPE’s and state-operator correspondence.
2. Setup
For definiteness, our computations will be carried out for Type IIA string theory
with intersecting D6-branes and Ω6-planes, whose world-volume spans the four non-
compact directions and a 3-cycle in some internal (Calabi-Yau) space. However, our
analysis can be easily converted to Type IIB framework by T-duality. In order to
ensure calculability we assume that the stacks of D6-branes supporting the Standard
Model intersect with each other in a region of the internal space that looks like a
factorizable torus T6 = T2 ×T2 ×T2 4.
1 3 3
In terms of three complex coordinates, with I = 1,2,3, we have
ZI = X2I+2 +iX2I+3 , Z∗ = X2I+2 −iX2I+3 (2.1)
I
In this framework, we consider two stacks of D6-branes which intersect at angles
θ = πa in the three tori, and without loss of generality we assume that
I I
1 > a ,a ,a > 0 . (2.2)
1 2 3
4Semi-realistic MSSM constructions on factorizable orbifolds can be found in [68,72–94].
– 4 –
X5 b X7 a X9
aa1b aa2b
ψ, ψ~,... _ ~_ a ψ_, ψ~_,... ψ_, ψ~_,... aa3b
ψ, ψ,... a
M ~
4 ψ, ψ,... ~
ψ, ψ,...
b b
X X X
4 6 8
Figure 3: At D-brane intersections live towers of states and their masses are multiples of
the intersection angle.
Supersymmetry follows if we also impose
a +a −a = 0 . (2.3)
1 2 3
In order to identify not only the massless states but also the lowest lying massive
states living at the intersection between two stacks of intersecting branes, we will
studythe1-looppartitionfunctionofstringsbeingstretchedbetweenthetwobranes.
3. Character valued partition function
The 1-loop partition function of open strings living at the intersection of two stacks
of branes is encoded in the vacuum amplitude on the annulus with one boundary
mapped onto one stack of N branes and the other onto another stack of M branes.
All the states transform according to the bi-fundamental representation (N,M) of
the gauge group U(N)×U(M). For simplicity we will not consider here un-oriented
strings that can appear in (anti)symmetric representations of unitary as well as
orthogonal or symplectic groups. These would require the inclusion of Ω-planes and
asaresultofthecorrespondingMo¨bius-stripamplitudes. Inordertokeeptrackofthe
helicity of each state, we consider the character-valued one-loop partition function
or helicity super-trace, that reads
(cid:81)
(cid:88) ϑ (v) ϑ (u )
Z = STr(zhqα(cid:48)m2) = |I|isinπv cGSO a I a I (3.1)
a ϑ (v)(cid:81) ϑ (u )
a 1 I 1 I
where q = exp2πiτ, z = exp2πiv, h denotes the helicity and |I| is the degeneracy
of the ground-state, i. e. number of intersections or of Landau levels (in the T-dual
case of magnetised branes). We also set u = a τ where τ = iT/2 for the annulus5.
I I
The multiplicity NM is understood. The GSO projection is implemented by the
sum over the spin structures with cGSO = −cGSO = −cGSO = 1 and cGSO = ±1.
3 4 2 1
The sign of I in combination with the sign of cGSO determines the chirality of the
1
5For the M¨obius-strip one would have τ =iT/2+1/2
– 5 –
massless fermion always present in the ground-state. The odd spin structure simply
contributes
(cid:81)
ϑ (v) ϑ (u ) 1
Z = ±|I|isinπv 1 I 1 I = ±|I|isinπv = ± |I|(z+1/2 −z−1/2) (3.2)
odd (cid:81)
ϑ (v) ϑ (u ) 2
1 I 1 I
Additional bosonic and fermionic zero-modes may appear depending on the values
of the “angles” θ = πa . After using Riemann identity, the contribution of the even
I I
spin structures becomes
(cid:32) (cid:33) (cid:32) (cid:33)
isinπv|I| (cid:88) 1 (cid:88) (cid:89) 1 (cid:88)
Z = ϑ (v +(cid:15) u ) ϑ (v −(cid:15) u )+(cid:15)u
even (cid:81) 1 I 1 I K
ϑ (v) ϑ (u ) 2 2
1 I 1 I (cid:15)=±1 I K I
(3.3)
(cid:80)
Assuming 1 (cid:29) a ≥ 0 as well as the “triangular inequality” a ≥ 2a , using
I I I K
Riemann identity for Jacobi ϑ functions, provided in appendix A, and neglecting all
(cid:80)
oscillators with ν = n, n+ (cid:15) a with n ≥ 1, we get only states with spin s ≤ 1
I I I
|I| (cid:110)
Zeven = −2(cid:81) (1−qaI) (z1/2 −z1q21(a1+a2+a3))(1−z−1/2q12(−a1+a2+a3)) (3.4)
I
(cid:111)
(1−z−1/2q21(a1−a2+a3))(1−z−1/2q12(a1+a2−a3))+(z → z−1)
|I| (cid:110) (cid:16)
= (cid:81) 2 q21(a1+a2−a3) +q12(a1−a2+a3) +q12(a1+a2−a3)
2 (1−qaI)
I
(cid:17)
+q12(3a1+a2+a3) +q21(a1+3a2+a3) +q12(a1+a2+3a3)
(cid:16) (cid:17)
−(1+qa1)(1+qa2)(1+qa3) z+1/2 +z−1/2
(cid:16) (cid:17) (cid:111)
+2qa1+a2+a3 z+1 +z−1 +...
2
Notice that the above formula is valid for any a > 0.
I
Supersymmetric case a +a −a = 0
1 2 3
Fordefiniteness,wehenceforthenforcethesupersymmetriccondition6 a +a −a = 0.
1 2 3
Using the expansion of the geometric series 1/(1−qaI) = 1+qaI +q2aI +q3aI +...
for small angles a (cid:28) 1 and assuming a = a + a > a ≥ a , the total partition
I 3 1 2 2 1
function becomes
Z = (1−z∓1/2)q0 +(2−z−1/2 −z1/2)(qa1 +qa2)+
all
+(2−z−1/2 −z1/2)(q2a1 +q2a2)+(z−1 −3z−1/2 +4−3z1/2 +z1)qa1+a2+
+(2−z−1/2 −z1/2)(q3a1 +q3a2)+(z−1 −4z−1/2 +6−4z1/2 +z1)(q2a1+a2 +qa1+2a2)+
+(2−z−1/2 −z1/2)(q4a1 +q4a2)+(z−1 −4z−1/2 +6−4z1/2 +z1)(q3a1+a2 +qa1+3a2)+
+(2z−1 −7z−1/2 +10−7z1/2 +2z1)q2a1+2a2 +...
(3.5)
6Any combination a ±a ±a =0 would preserve some supersymmetry.
1 2 3
– 6 –
Therefore, the states living at the intersection of the two branes are organized in
multiplets with different masses:
• Massless sector (from q0)
Zmassless = |I|(1−z±1/2) (3.6)
chiral
produces a massless (anti-)chiral multiplet.
• Massive states with α(cid:48)m2 = ka (from qka1,2 with k = 1,2,...)
1,2
(cid:16) (cid:17)
Zα(cid:48)m2=ka1,2 = |I| 2−z−1/2 −z1/2 qka1,2 (3.7)
form massive chiral multiplets (one chiral and one anti-chiral, as we will see).
• Massive states with α(cid:48)m2 = ka +a = a +(k−1)a (from qka1+a2 with k =
1 2 3 1
1,2,...)
(cid:110)(cid:16) (cid:17) (cid:16) (cid:17)(cid:111)
Zα(cid:48)m2=a3+(k−1)a1 = |I| 2−z−1/2 −z1/2 + 2−2z−1/2 −2z1/2 +z−1 +z1 qka1+a2
(3.8)
form a massive chiral and a massive vector multiplet. The same applies to
states with α(cid:48)m2 = a +ka = a +(k−1)a .
1 2 3 2
• Massive states with α(cid:48)m2 = 2a = 2a +2a (from q2a1+2a2)
3 1 2
(cid:110) (cid:16) (cid:17) (cid:16) (cid:17)(cid:111)
Zα(cid:48)m2=2a3 = |I| 3 2−z−1/2 −z1/2 +2 2−2z−1/2 −2z1/2 +z−1 +z1 q2a3
(3.9)
form three massive chiral and two massive vector multiplets.
The precise identification of the lowest lying physical states allows us to deter-
mine their vertex operators: a task we will accomplish in the next section.
4. States and vertex operators
Let us write down the vertex operators (VO’s) for the massless and lowest-lying
massive states with α(cid:48)m2 = a with I = 1,2,3, starting from the massless ones.
I
4.1 Gauge sector
Recall that the Left-Handed (LH) and Right-Handed (RH) gaugino vertex operators
in the canonical −1/2 super-ghost picture read
3 3
(cid:89) (cid:89)
V = λα(k)S eiϕI/2e−ϕ/2eipX , V = λ˜α˙(k)C e−iϕI/2e−ϕ/2eipX (4.1)
λ α λ˜ α˙
I=1 I=1
– 7 –
where S and C denote SO(1,3) spin fields of opposite chirality. They transform in
α α˙
theadjointrepresentationNN+MMofthegaugegroupanddeterminetheunbroken
SUSY charges
(cid:73) 3 (cid:73) 3
dz (cid:89) dz (cid:89)
Q(−1/2) = S eiϕI/2e−ϕ/2 , Q˜(−1/2) = C e−iϕI/2e−ϕ/2 (4.2)
α 2πi α α˙ 2πi α˙
I=1 I=1
that should act “locally” on all physical vertex operators (only integer powers in the
OPE). As a result the three ‘internal’ U(1) R-charges r are quantised. For instance
I
(cid:80)
r = −1/2 for massless LH matter fermions in the canonical super ghost picture,
I I
as we will see momentarily. For later use, let us recall the form of the SUSY charges
in the +1/2 super-ghost picture
(cid:73) 3
dz (cid:16) (cid:17)(cid:89)
Q(+1/2) = e+ϕ/2 S ∂ZIψ∗ +∂X σµ Cα˙ eiϕI/2 (4.3)
α 2πi α I µ αα˙
I=1
(cid:73) 3
dz (cid:16) (cid:17)(cid:89)
Q˜(+1/2) = e+ϕ/2 C ∂Z∗ψI +∂X Sασµ e−iϕI/2 (4.4)
α˙ 2πi α˙ I µ αα˙
I=1
4.2 Massless sector
For a choice of GSO and sign of I, the massless chiral multiplets, living at the |I|
intersections and transforming in the (N,M) representation, consist in one scalar
and a LH fermion each, whose internal states read
φ(k) : ψ3 |a ,a ,a (cid:105) , χα(k) : |a ,a ,a (cid:105) (4.5)
−12+a3 1 2 3 NS 1 2 3 R
where7 |a ,a ,a (cid:105) = lim σ σ σ† (z)|0(cid:105) and vertex operators (VO’s hence-
1 2 3 NS z→0 a1 a2 a3 NS
forth) 8
V0 = φ(k)σ σ σ† ei[a1ϕ1+a2ϕ2+(1−a3)ϕ3]e−ϕeikX (4.6)
φ a1 a2 a3
Vχ0 = χα(k)Sασa1σa2σa†3ei[(a1−12)ϕ1+(a2−21)ϕ2+(12−a3)ϕ3]e−ϕ/2eikX (4.7)
The GSO projection determines the chirality of the four-dimensional spinor in terms
of the U(1) world-sheet charge of the respective vertex operator. The latter is given
by (cid:80)3 r , where the r are the coefficients of ϕ , that bosonize the internal world-
I=1 I I I
sheet fermions, i. e. ψI = exp(iϕ ), ψ∗ = exp(−iϕ ). If the sum adds up to
I I I
−1/2 (mod 2) the spinor is chiral (LH), while if the sum is +1/2 (mod 2) the spinor
is anti-chiral (RH). Similarly complex bosons in chiral multiplets, pairing with LH
7Our notation here differs from the one in [62] since |a ,a ,a (cid:105)here ↔ |a ,a ,−a (cid:105)there. One
1 2 3 NS 1 2 3 NS
also has σ† =σ .
a 1−a
8Foradetaileddiscussiononvertexoperatorsofmasslessstatesforarbitraryintersectionangles,
see[95,96],forageneralizationtomassivestates,see[62],andforadiscussiononinstantonicstates
at the intersection of D-instanton and D-brane at arbitrary angles, see [97].
– 8 –
fermions, carry (cid:80)3 r = +1, while those in anti-chiral multiplets, pairing with RH
I=1 I
fermions, carry (cid:80)3 r = −1. In a not-so-lose sense this charge can be identified
I=1 I
with the R-charge of the N = 1 SUSY algebra in that – in the canonical super- ghost
picture – the super-charges and the gaugini carry R-charges ±3/2. Changing picture
vertex operators and super-charges have no definite R-charge, which is not an exact
continuous symmetry in fact.
The BRST charge is (focusing on the matter part)
(cid:73) (cid:26) (cid:18) 3 3 (cid:19)
dz η (cid:88) (cid:88)
Q = eφ√ i∂Xµψ + ∂ZI ψ∗ + ∂Z∗ψI (4.8)
BRST 2πi 2α(cid:48) µ I I
I=1 I=1
c (cid:20) α(cid:48) (cid:88)3 (cid:16) α(cid:48) α(cid:48) (cid:17)(cid:21)(cid:27)
+ i∂Xµi∂X − ψµ∂ψ − ∂ZI∂Z∗ + ψI∂ψ∗ + ψ∗∂ψI +...
α(cid:48) µ 2 µ I 2 I 2 I
I=1
and the physical state condition is
[Q ,V] = 0 (4.9)
BRST
and using the OPE’s we get a simple and a double pole.
• The simple pole identically vanishes for the scalar. For the fermions, the van-
ishing of the simple pole requires the Dirac equation of motion. In this massless
case it is just
kµσ¯α˙αχ (k) = 0 (4.10)
µ α
• Vanishing of the double pole always yields the standard mass-shell condition
k2 = 0 (4.11)
for any massless bosonic or fermionic VO. For massive ones one gets
p2 +m2 = 0 (4.12)
Henceforth we will use k to denote light-like momenta and p to denote time-like
ones. Dependence on the momenta of the ‘polarisations’ or ‘wave-functions’ of
the vertex operators will be explicitly shown as (cid:96)(k) or H(p) for convenience.
4.3 Massive multiplets
Here we study separately the α(cid:48)m2 = a and the α(cid:48)m2 = a cases. LH fermions in
1,2 3
chiral multiplets in the (N,M¯ ) representation will be denoted by χI(p), their scalar
f
partners by φI(p) with I = 1,2,3 referring to their mass α(cid:48)m2 = a and f = 1,...n
f I I
their multiplicity. LH fermions and scalars in chiral multiplets in the (M,N¯) will be
denoted by χ˜I(p), their scalar partners by φ˜I(p). States in anti-chiral multiplets will
f f
be distinguished by a dagger.
– 9 –