For our splitting and coupling constant ˆ =1+i2, the relationship (3.34) remains the same,
↵1(t) = 1(t) 1(t), ↵2(t) = 1(t) 2(t), (5.7) although the coefficients now explicitly carry time dependence, and we set
2 = 0 due to reality of the coupling. Then, we defineT from1 by following the exact same procedure as in section 3.2. The higher order expansions of the fields are more involved than the static analysis. They were carried out by a program published by Strydom in [67], which we needed to adapt for our purposes.
The reason for adapting the code lies within the gauge fixing in this model. In holography, we commonly choose the radial gauge for the gauge field,
az = 0, (5.8)
as this is a natural choice due to the radial direction being the emergent holographic direction. Thus, we are left with the temporal component at
and its equation of motion, which is thet-component of (5.3). This fixes the gauge only up to a function oft, as explained in more detail in appendix C.
In this thesis, we do not utilise this residual gauge freedom to fix the phase of the scalar asymptotically, as was done in [67]. There, the phase of the scalar was chosen to vanish asymptotically, which yields 2 = 0. It was found that, due to the asymptotics of the gauge field in AdS2, obtaining this result requires a large gauge transformation which alters the physics. Since we want to compare e.g. with the results of [68], in which such a large gauge transformation was not applied, we leave the scalar’s phase free throughout the bulk.
The model as described above can easily be reduced to the static model of [60], described in section 3.2, by setting any occurring temporal derivatives to zero. Especially, the model still features a phase transition at a critical tem-perature Tc ⇡ TK which transfers to a critical coupling c = 1/log(TK/Tc), whereTK denotes the Kondo temperature. For our choice of the representa-tion, C = 1/2, the numerical value for the critical coupling is given by
c ⇡8.9796, (5.9)
which, of course, is the same number as in (3.46).
(i) Compute the boundary expansions of the fields at the event horizon and asymptotic infinity up to some convenient order,
(ii) restrict the computational domain to the region Rnum = ["bnd,"hor] between two cuto↵ parameters at the event horizon and asymptotic infinity, and
(iii) compute the fields in Rnum numerically by providing boundary con-ditions at "bnd and "hor which originate from an approximation of the fields by applying the respective boundary expansions up to the cuto↵s.
In the case at hand, however, the procedure is surprisingly involved. The rea-son is that we need to fix the scalar field’s mass to saturate the Breitenlohner-Freedman bound in order to obtain the correct scaling dimension for the dual operator as discussed in 3.2.2. This induces many logarithmic terms in the boundary expansion, as we already encountered in the static backreaction in chapter 4. In the time dependent setting presented in this chapter, even more logarithmic terms can be turned on and solving for all the coefficients becomes tedious.
So, especially step (i) becomes much harder. Fortunately, the algorithm introduced in [67] is capable to find the coefficients order by order in an expansion around the asymptotic boundary. However, finding the right cuto↵
values in (ii) and evolving the fields by supplying approximate boundary values from the expansions in (iii) is tricky, because it is numerically hard to distinguish between terms which only di↵er in their logarithmic power.
In order to address these issues, we replace the dependent variables ak
and i, wherekdenotes the left-over component of the gauge field after fixing the gauge, by substituted fields ˜ak and ˜i. Those are built by subtracting the boundary terms of the respective fields up to a convenient order:
˜
ak ⌘ak sa, (5.10)
˜i ⌘ i s ,i, (5.11)
where the subtracted termssa and s ,i are given by sa= Q
z +µ(t) +O z1 , (5.12)
s ,i=p
z(↵i(t) log(z) + i(t)) +O z3/2 , (5.13) and theO symbols denote higher order terms. The substituted fields ˜ak and
˜i are replacing the original fields in the equations of motion. This way, we
transfer the dependency of the system on the involved boundary coefficients µ(t),↵i(t) and i(t) from the fields themselves to the equations of motion.
Moreover, with a static event horizon due to the probe limit, there is no need to use the cuto↵ at the horizon and appropriate horizon conditions.
Instead, we choose our discretisation scheme in the radial direction as an expansion of the fields in Chebyshev polynomials and choose a point inside the event horizon as the boundary of the computational domain Rnum. As Chebyshev polynomials build a complete set of regular functions and the computational domain includes the event horizon, we implicitly require reg-ularity at the horizon. The metric has the feature that the future lightcone tilts into the event horizon once we cross it. This way, points in the event horizon cannot a↵ect points outside and ingoing boundary conditions are satisfied automatically.
The temporal direction is discretised on an evenly distributed grid, which sets bounds on the temporal resolution. Especially, as we did not implement an adaptive temporal grid, we need to know which frequencies have to be resolved before we start the run. Without a semi-analytic analysis of the quasinormal modes, this can only be done by trial and error. We perform perturbative Gaussian quenches over a fixed background value for the Kondo coupling 0, given by
T(t) = 0 1 +a exp s2(t t0)2 , (5.14) where v0 is the time around which the quench is centred, the parameter s is describing the steepness of the quench, and a is the ratio of maximal am-plitude to background amam-plitude of the coupling. Quenches like this kick the system slightly out of equilibrium and we can find the respective quasi-normal mode at 0 by fitting the ring-down of the fields to a quasinormal behaviour. In addition to that, quenches starting in the normal phase feature another quasinormal-like behaviour at early times. Here, the perturbation is exponentially growing in time, with a time constant which can be read o↵ an-alytically from performing perturbation theory around the normal phase, in which the scalar vanishes. Finally, the steepness parameter s of the quench itself is setting a temporal scale. In order to set the temporal resolution properly, we must ensure that the smallest time scale from above is resolved by the temporal grid.
The time evolution scheme is chosen to be Crank-Nicolson-like, as al-ready described in [67]. For some more details about the numerical evolution scheme, we refer the reader to appendix C.