Abstract
The equilibrium optical phonons of graphene are well characterized in terms of anharmonicity and electron–phonon interactions; however, their nonequilibrium properties in the presence of hot charge carriers are still not fully explored. Here we study the Raman spectrum of graphene under ultrafast laser excitation with 3 ps pulses, which trade off between impulsive stimulation and spectral resolution. We localize energy into hot carriers, generating nonequilibrium temperatures in the ~1700–3100 K range, far exceeding that of the phonon bath, while simultaneously detecting the Raman response. The linewidths of both G and 2D peaks show an increase as function of the electronic temperature. We explain this as a result of the Dirac cones’ broadening and electron–phonon scattering in the highly excited transient regime, important for the emerging field of graphenebased photonics and optoelectronics.
Introduction
The distribution of charge carriers has a pivotal role in determining fundamental features of condensed matter systems, such as mobility, electrical conductivity, spinrelated effects, transport, and optical properties. Understanding how these proprieties can be affected and, ultimately, manipulated by external perturbations is important for technological applications in diverse areas, ranging from electronics to spintronics, optoelectronics and photonics^{1,2,3}.
The current picture of ultrafast light interaction with singlelayer graphene (SLG) can be summarized as follows^{4}. Absorbed photons create optically excited electron–hole (e–h) pairs. The subsequent relaxation towards thermal equilibrium occurs in three steps. Ultrafast electron–electron (e–e) scattering generates a hot Fermi–Dirac distribution within the first tens fs^{5}. The distribution then relaxes due to scattering with optical phonons (ph; electron–phonon (e–ph) coupling), equilibrating within a few hundred fs^{6,7}. Finally, anharmonic decay into acoustic modes establishes thermodynamic equilibrium on the ps timescale^{8,9,10}.
Raman spectroscopy is one of the most used characterization techniques in carbon science and technology^{11}. The measurement of the Raman spectrum of graphene^{12} triggered a substantial effort to understand phonons, e–ph, magneto–ph, and e–e interactions in graphene, as well as the influence of the number and orientation of layers, electric or magnetic fields, strain, doping, disorder, quality and types of edges, and functional groups^{13,14,15}. The Raman spectra of SLG and few layer graphene (FLG) consist of two fundamentally different sets of peaks. Those, such as D, G, 2D, present also in SLG, and due to inplane vibrations^{12}, and others, such as the shear (C) modes^{16} and the layer breathing modes^{17,18} due to relative motions of the planes themselves, either perpendicular or parallel to their normal. The G peak corresponds to the high frequency E_{2g} phonon at Γ. The D peak is due to the breathing modes of sixatom rings, and requires a defect for its activation^{19,20,21}. It originates from transverse optical (TO) phonons around the Brillouin Zone edge K[19], it is active by double resonance (DR)[20] and, due to a Kohn Anomaly at K[22], it is dispersive with excitation energy. The 2D peak is the D overtone and originates from a process where momentum conservation is fulfilled by two phonons with opposite wavevectors. It is always present since no defects are required for this process^{12}.
Raman spectroscopy is usually performed under continuous wave (CW) excitation, therefore probing samples in thermodynamic equilibrium. The fast e–e and e–ph nonradiative recombination channels establish equilibrium conditions between charge carriers and lattice, preventing the study of the vibrational response in presence of an hot e–h population. Using an average power comparable to CW illumination (a few mW), ultrafast optical excitation can provide large fluences (~1−15 J m^{−2} at MHz repetition rates) over sufficiently short timescales (0.1–10 ps) to impulsively generate a strongly outofequilibrium distributions of hot e–h pairs^{4,8,23,24}. The potential implications of coupled e and ph dynamics for optoelectronics were discussed for nanoelectronic devices based on CW excitation^{25,26,27,28,29}. However, understanding the impact of transient photoinduced carrier temperatures on the colder SLG ph bath is important for mastering out of equilibrium e–ph scattering, critical for photonics applications driven by carrier relaxation, such as ultrafast lasers^{30}, detectors^{1,3} and modulators^{31}. E.g, SLG can be used as a much broaderband alternative to semiconductors saturable absorbers^{30}, for modelocking and Q switching^{1,30}.
Here we characterize the SLG optical ph at high electronic temperature, T_{e}, by performing Raman spectroscopy under pulsed excitation. We use a 3 ps pulse to achieve a tradeoff between the narrow excitation bandwidth required for spectral resolution \(\left( {\frac{{\delta \nu }}{c} \le } \right.\) 10 cm^{−1}, being v[Hz] the laser frequency and c the speed of light, a condition met under CW excitation) and a pulse duration, δt, sufficiently short (δt ≤ 10 ps, achieved using ultrafast laser sources) to generate an highly excited carrier distribution over the equilibrium ph population, being those two quantities Fourier conjugates^{32} \(\left( {\frac{{\delta \nu }}{{\delta t}} \le 14.7{\mathrm{cm}}^{  1}\mathrm{ps}} \right)\). This allows us to determine the dependence of both ph frequency and dephasing time on T_{e}, which we explain by a broadening of the Dirac cones.
Results
Hot photoluminescence
Figure 1a plots a sequence of antiStokes (AS) Raman spectra of SLG following ultrafast excitation at 1.58 eV, as a function of excitation power P_{L}. The broad background stems from hot photoluminescence (PL) due to the inhibition of a full nonradiative recombination under high excitation densities^{8,26,33,34}. This process, absent under CW excitation in pristine SLG^{35}, is due to ultrafast photogeneration of charge carriers in the conduction band, congesting the e–ph decay pathway, which becomes progressively less efficient with increasing fluence. This nonequilibrium PL recalls the gray body emission and can be in first approximation described by Planck’s law:^{8,26,29,33}
where η is the emissivity, defined as the dimensionless ratio of the thermal radiation of the material to the radiation from an ideal black surface at the same temperature as given by the Stefan–Boltzmann law^{36}, τ_{em} is the emission time and \({\cal R}(\hbar \omega )\) is the frequency dependent, dimensionless responsivity of our detection chain. Refs^{8,29,33} reported that, although Eq. 1 does not perfectly reproduce the entire gray body emission, the good agreement on a ~0.5 eV energy window is sufficient to extract T_{e}. By fitting the background of the Raman spectra with Eq. 1 (solid lines in Fig. 1a) we obtain T_{e} as a function of P_{L}. Figure 1b shows that T_{e} can reach up to 3100 K under our pulsed excitation conditions.
An upper estimate for the lattice temperature, T_{1}, can be derived assuming a full thermalization of the optical energy between vibrational and electronic degrees of freedom, i.e., evaluating the local equilibrium temperature, T_{eq}, by a specific heat argument (see Methods). We get T_{eq}(P_{max}) ~ 680 K at the maximum excitation power, P_{max} = 13.5 mW. This is well below the corresponding T_{e}, indicating an outofequilibrium distribution of charge carriers. Thus, over our 3 ps observation timescale, T_{1} is well below T_{eq}.
Out of equilibrium Raman response
Figure 1c plots the AS and S G peaks, together with fits by Lorentzians (blue lines) convoluted with the laser bandwidth (~9.5 cm^{−1}) and spectrometer resolution (~6 cm^{−1}), which determine the instrumental response function, IRF (see Methods). The S data have a larger noise due to a more critical background subtraction, which also requires a wider accessible spectral range (see Methods). For this reason, we will focus on the AS spectral region, with an higher spectrometer resolution (1.2 cm^{−1}), Fig. 2. We obtain a full width at half maximum of the G peak, FWHM(G) ~21 cm^{−1}, larger than the CW one (~12.7 cm^{−1}). Similarly, we get FWHM(2D) ~50–60 cm^{−1} over our P_{L} range, instead of FWHM(2D) ~29 cm^{−1} as measured on the same sample under CW excitation. To understand the origin of such large FWHM(G) andFWHM(2D) in pulsed excitation, we first consider the excitation power dependence of the SLG Raman response in the 1.53–13.5 mW range (the lower bound is defined by the detection capability of our setup). This shows that the position of the G peak, Pos(G), is significantly blueshifted (as reported for graphite in ref^{23}.), while the position of the 2D peak, Pos(2D), is close to that measured under CW excitation, and both FWHM(G) and FWHM(2D) increase with P_{L}. Performing the same experiment on Si proves that the observed peaks broadening is not limited by our IRF (see Methods). Moreover, even the low resolution S data of the G band, collected in the range 1.8–7.0 mW (a selection is shown in Fig. 1c), display a broadening ((8 ± 4)10^{−3} cm^{−1} K^{−1}) and upshift ((2.8 ± 1.8)10^{−3} cm^{−1} K^{−1}), compatible with that of the highresolution AS measurements (Fig. 3d, e) (7.4 ± 0.5)10^{−3} cm^{−1} K^{−1} and (3.2 ± 0.2)10^{−3} cm^{−1} K^{−1}.
We note that phonons temperature estimates based on the AS/S intensity ratio^{37,38} are not possible in our case due to two concurring effects. First, SLG’s resonant response to any optical wavelength gives a non trivial wavelength dependent Raman excitation profile, which modifies the Raman intensities with respect to the nonresonant case. Consequently, the AS/S ratio is no longer straightforwardly related to the thermal occupation^{39}. Second, in SLG a S ph may be subsequently annihilated by a correlated AS event. This may result into an extra AS pumping which does not allow to relate AS/S ratio and ph temperature via the thermal occupation factor^{40}. Accordingly, the AS/S ratio approaching one at the largest excitation power in Fig. 1c (black circles) does not necessary imply a large increase of the G phonon temperature.
Discussion
Figure 3 plots Pos(2D), FWHM(2D), Pos(G), FWHM(G) as a function of T_{e}, estimated from the hotPL. A comparison with CW measurements (633 nm) at room temperature (RT) is also shown (blue diamonds). Under thermodynamic equilibrium, the temperature dependence of the Raman spectrum of SLG is dominated by anharmonicity, which is responsible for mode softening, leading to a redshift of the Raman peaks^{10,41,42}. This differs from our experiments (Fig. 4a–d), in which Pos(G) has an opposite trend (blue shift), and Pos(2D) is nearly T_{e} independent, in agreement with density functional perturbation theory (DFPT) calculations, giving a variation of the 2D peak ΔPos(2D) ~5 cm^{−1} in the range T_{e} = 300–3000 K (see Methods). This indicates the lack of significant anharmonic effects and suggests a dominant role of e–ph interaction on FWHM(G) and Pos(G), in the presence of a cold phonon bath at constant T_{1} decoupled from the (large) T_{e}.
To derive the temperature dependence of such parameters, we first compute the phonon selfenergy \({\Pi}(q = 0,\omega _{\mathrm{G}}^0)\), as for refs.^{22,43,44},
Here \(\xi = g^2/(2\hbar m_{\mathrm{a}}\omega _{\mathrm{G}}^0v_{\mathrm{F}}^2) = 4.43 \times 10^{  3}\) is a dimensionless constant, v_{F} is the Fermi velocity, \(\tilde \varepsilon\) is the upper cutoff of the linear dispersion ε = v_{F}k, m_{a} is the carbon atom mass, \(\hbar \omega _{\mathrm{G}}^0 = 0.196\)eV the bare phonon energy, δ is a positive arbitrary small number (< 4 meV), g ~ 12.3 eV is proportional to the e–ph coupling^{6,22,43,45}, z, z′ are the energy integration variables and f_{F}(z − E_{F}) is the Fermi–Dirac distribution with E_{F} the Fermi energy. Although our samples are doped, E_{F} significantly decreases as a function of T_{e}^{25}. Hence, we assume E_{F} = 0 in the following calculations. The two indexes s,s′ = \(\mp 1\) denote the e and h branches, and M_{s}(z, ε) is the corresponding spectral function, which describes the electronic dispersion.
The selfenergy expressed by Eq. 2 renormalizes the phonon Green’s function according to the Dyson’s equation:^{46}
so that ΔPos(G) and FWHM(G) can be written as:
where h is the Planck constant. FWHM(G) can be further simplified since the evaluation of \({\mathrm{Im}}{\Pi}(0,\omega _{\mathrm{G}}^0,T_{\mathrm{e}})\) leads to \(\delta (z  z{\prime}  \hbar \omega _{\mathrm{G}}^0)\) in Eq. 2, so that we get:
In the limit of vanishing broadening of the quasiparticle state, the SLG gapless linear dispersion is represented by the following spectral function:^{46}
This rules the energy conservation in Eq. 5 and allows only transitions between h and e states with energy difference \(2\varepsilon = \hbar \omega _{\mathrm{G}}^0\). Thus, we get:^{22,43,44}
where \({{\mathrm{FWHM}}}{({\mathrm{G}})}^0 = \frac{{\pi \xi \hbar \omega _{\mathrm{G}}^0}}{{2hc}} \sim 11\)cm^{−1} ^{10}. This value, with the additional ~2 cm^{−1} contribution arising from anharmonic effects^{10}, is in agreement with the CW measurements at T_{e} = T_{eq} = 300 K (see diamond in Fig. 3e), corresponding to fluences ≪1 J m^{−2}. Eq. 7 also shows that, as T_{ e } increases, the conduction band becomes increasingly populated, making the phonon decay channel related to e–h formation progressively less efficient and leading to an increase of the ph decay time (Fig. 4b). This leads to a decrease of FWHM(G) for increasing T_{e} (black solid line in Fig. 3e), in contrast with the experimentally observed increase (blue circles in Fig. 3e).
A more realistic description may be obtained by accounting for the effect of T_{e} on the energy broadening (γ_{e}) of the linear dispersion M_{s}(z,ε), along with the smearing of the Fermi function. γ_{e}(z,T_{e}) can be expressed, to a first approximation, as the sum of three terms:^{47}
where γ_{ee}, γ_{ep} and γ_{def} are the e–e, e–ph and defect contributions to γ_{e}. The only term that significantly depends on T_{e} is γ_{ee}, while the others depend on the energy z^{10,44,47,48,49,50}.
The linear dependence of γ_{ee} on T_{e}^{51} can be estimated from its impact on FWHM(2D). The variation of FWHM(2D) with respect to RT can be written as:^{42}
where \([\partial {{\mathrm{POS}}(2{\mathrm{D}})}/\partial (h\nu _{{\mathrm{laser}}})]/2 = \frac{1}{{ch}}v_{\mathrm{ph}}/v_{\mathrm{F}}{\mathrm{\sim }}100\)cm^{−1} eV^{−1}^{13,52}, i.e., the ratio between the ph and Fermi velocity, defined as the slope of the phononic (electronic) dispersion at the ph (e) momentum corresponding to a given excitation laser energy hν_{laser}^{13}. Since the DR process responsible for the 2D peak involves the creation of e–h pairs at energy \(\mp\) hν_{laser}/2, the change of FWHM(2D) allows us to estimate the variation of γ_{e} at \(z = h\nu _{{\mathrm{laser}}}/2 \simeq 0.8\)eV. Then, γ_{ep} and γ_{def}, both proportional to z (γ_{ep},γ_{def} ∝ z), will give an additional constant contribution to FWHM(2D), but not to its variation with T_{e}. Our data support the predicted^{51} linear increase of γ_{ee} with T_{e}, with a dimensionless experimental slope \(\alpha _e \simeq 0.73\), Fig. 3c.
In order to compute FWHM(G) from Eq. 2, we note that the terms γ_{ep} and γ_{def} are negligible at the relevant low energy \(z = \hbar \omega _{\mathrm{G}}/2 \sim 0.1\)eV \(\ll h\nu _{{\mathrm{laser}}}/2\). Hence \(\gamma _{\mathrm{e}}(z,T_{\mathrm{e}}) \simeq \gamma _{{\mathrm{ee}}}(T_{\mathrm{e}})\).
The Dirac cone broadening can now be introduced by accounting for γ_{ e } in the spectral function of Eq. 6:
accordingly, all the processes where the energy difference sε(k) − s′ε(k′) + ħω_{0} is less than 2γ_{e} (which guarantees the overlap between the spectral functions of the quasiparticles) will now contribute in Eq. 2. Amongst them, those transitions within the same (valence or conduction) band, as shown in Fig. 4d.
The broadened interband contributions still follow, approximately, Eq. 7 (Fig. 4e). However, the Dirac cone broadening gives additional channels for G phonon annihilation by carrier excitation. In particular, intraband transitions within the Dirac cone are now progressively enabled for increasing T_{e}, as sketched in Fig. 4d. In Fig. 4e the corresponding contributions to FWHM(G) are shown. Calculations in the weakcoupling limit^{51} suggest that γ_{e}(T_{e}) should be suppressed as z → 0, due to phase–space restriction of the Dirac cone dispersion. Our results, however, indicate that this effect should appear at an energy scale smaller than ℏω_{G}/2, as the theory captures the main experimental trends, just based on a zindependent γ_{e}(T_{e}).
Critically, the G peak broadening has a different origin from the equilibrium case^{53}. The absence of anharmonicity would imply a FWHM(G) decrease with temperature due to the e–ph mechanism. However, the Dirac cone broadening reverses this trend into a linewidth broadening above T_{e} = 1000K producing, in turn, a dephasing time reduction, corresponding to the experimentally observed FWHM(G) increase. The blue shift of the G peak with temperature is captured by the standard e–ph interaction, beyond possible calibration accuracy. Importantly, the Dirac cone broadening does not significantly affect Pos(G).
In conclusion, we measured the Raman spectrum of SLG with impulsive excitation, in the presence of a distribution of hot charge carriers. Our excitation bandwidth enables us to combine frequency resolution, required to observe the Raman spectra, with short pulse duration, needed to create a significant population of hot carriers. We show that, under these strongly nonequilibrium conditions, the Raman spectrum of graphene cannot be understood based on the standard low fluence picture, and we provide the experimental demonstration of a broadening of the electronic linear dispersion induced by the highly excited carriers. Our results shed light on a peculiar regime of nonequilibrium Raman response, whereby the e–ph interaction is enhanced. This has implications for the understanding of transient charge carrier mobility under photoexcitation, important to study SLGbased optoelectronic and photonic devices^{27,28}, such as broadband light emitters^{29}, transistors and optical gain media^{54}.
Methods
Sample preparation and CW raman characterization
SLG is grown on a 35 μm Cu foil, following the process described in Refs^{55,56}. The substrate is heated to 1000 °C and annealed in hydrogen (H_{2}, 20 sccm) for 30 min. Then, 5 sccm of methane (CH_{4}) is let into the chamber for the following 30 min so that the growth can take place^{55,56}. The sample is then cooled back to RT in vacuum (∼1 mTorr) and unloaded from the chamber. The sample is characterized by CW Raman spectroscopy using a Renishaw inVia Spectrometer equipped with a 100× objective. The Raman spectrum measured at 514 nm is shown in Fig. 5 (red curve). This is obtained by removing the nonflat background Cu PL^{57}. The absence of a significant D peak implies negligible defects^{12,13,21,58}. The 2D peak is a single sharp Lorentzian with FWHM(2D) ∼23 cm^{−1}, a signature of SLG^{12}. Pos(G) is ∼1587 cm^{−1}, with FWHM(G) ∼14 cm^{−1}. Pos(2D) is ∼2705 cm^{−1}, while the 2D to G peak area ratio is ∼4.3. SLG is then transferred on glass by a wet method^{59}. Polymethyl methacrylate (PMMA) is spin coated on the substrate, which is then placed in a solution of ammonium persulfate (APS) and deionized water. Cu is etched^{55,59}, the PMMA membrane with attached SLG is then moved to a beaker with deionized water to remove APS residuals. The membrane is subsequently lifted with the target substrate. After drying, PMMA is removed in acetone leaving SLG on glass. The SLG quality is also monitored after transfer. The Raman spectrum of the substrate shows features in the D and G peak range, convoluted with the spectrum of SLG on glass (blue curve in Fig. 5). A pointtopoint subtraction is needed to reveal the SLG features. After transfer, the D peak is still negligible, demonstrating that no significant additional defects are induced by the transfer process. The fitted Raman parameters indicate p doping ∼250 meV^{50,60}.
Before and after the pulsed laser experiments, equilibrium CW measurements are performed at RT using a microRaman setup (LabRAM Infinity). A different energy and momentum of the D phonon is involved, for a given excitation wavelength, in the S or AS processes, due to the phonon dispersion in the DR mechanism^{61,62}. Hence, in order to measure the same D phonon in S and AS, different laser excitations (ν_{laser}) must be used according to \(\nu_{\mathrm{laser}}^{\rm S} = \nu _{{\rm laser}}^{{\rm AS}} + {\rm cPos}(2{\mathrm{D}})\)^{13,63,64}. Given our pulsed laser wavelength (783 nm), the corresponding CW excitation would be ∼649.5 nm. Hence, we use a 632.8 nm HeNe source, accounting for the small residual wavelength mismatch by scaling the phonon frequency as \(\frac{{\mathrm{dPos}}(2{\mathrm{D}})}{{\mathrm{d}\nu}_{\mathrm{laser}}} = 0.0132/c\) ^{13}.
Pulsed raman measurements
The psRaman apparatus is based on a modelocked Er:fiber amplified laser at ∼1550 nm, producing 90fs pulses at a repetition rate RR = 40 MHz. Using secondharmonic generation in a 1 cm periodically poled Lithium Niobate crystal^{65}, we obtain 3 ps pulses at 783 nm with a ∼9.5 cm^{−1} bandwidth. The beam is focused on SLG through a slightly underfilled 20× objective (NA = 0.4), resulting in a focal diameter D = 5.7 μm. Backscattered light is collected by the same objective, separated with a dichroic filter from the incident beam and sent to a spectrometer (with a resolution ~0.028 nm/pixel corresponding to ~1.2 cm^{−1}). The overall IRF, therefore, is dominated by the additional contribution induced by the finite excitation pulse bandwidth. Hence, in order to extract the FWHM of the Raman peaks, our data are fitted convolving a Lorentzian with the spectral profile of the laser excitation.
When using ultrafast pulses, a nonlinear PL is seen in SLG^{8}. Such an effect is particularly intense for the S spectral range^{34,66}. The S signal in Fig. 1c is obtained as the difference spectrum of two measurements with excitation frequencies offset by ∼130 cm^{−1}, resulting in PL suppression. The background subtraction requires in this case a wider spectral range, at the expenses of spectrometer resolution which is reduced to ~0.13 nm/pixel, corresponding to ∼6 cm^{−1}, as additional contribution to the IRF. Although this procedure allows to isolate the S Raman peaks, the resulting noise level is worse than for AS. For this reason we mostly focus on the AS features.
To verify that the observed peaks broadening is not limited by our IRF, we perform the same experiment on a Si substrate (Fig. 6a). For this we retrieve, after deconvolution of the IRF, the same Raman linewidth measured in the CW excitation regime (Fig. 6a). The FWHM of the Si optical phonon is independent of P_{L}, in contrast with the welldefined dependence on P_{L} observed in SLG, Fig. 6b.
Estimate of T_{eq}
Photoexcitation of SLG induces an excess of energy in the form of heat per unit area, Q, that can be expressed as:
where A = 2.3% is the SLG absorption, approximated to the undoped case^{67}, \(W \sim 2.8{\mu}{\mathrm{m}}\) is the waist of focused beam and RR = 40 MHz is the repetition rate of the excitation laser. The induced T_{eq} can be derived based on two assumptions: (i) in our ps timescale the energy absorbed in the focal region does not diffuse laterally, (ii) the energy is equally distributed on each degree of freedom (electrons, optical and acoustic ph). Then, Q can be described as:
where C(T) is the SLG Tdependent specific heat. In the 300–700 K range, C(T) can be described as:^{68} C(T) = aT + b, where a = 1.35×10^{−6}J K^{−2} m^{−2} and b = 1.35×10^{−4}J K^{−1} m^{−2}. Therefore, considering Eqs. 11,12, for P_{L} = P^{max} = 13.5 mW, we get \(T_{{\mathrm{eq}}} \sim 680\)K, well below the corresponding T_{e}, indicating an outofequilibrium condition (T_{l} < T_{eq} < T_{e}). Any contributions from the substrate and taking into account for the heat profile would contribute in reducing even further T_{l} estimation.
Estimate of Pos(2D) as a function of T_{e}
We perform calculations within the local density approximation in DFPT^{69,70}. We use the experimental lattice parameter 2.46 Å^{71} and plane waves (45 Ry cutoff), within a normconserving pseudopotential approach^{70}. The electronic levels are occupied with a finite fictitious T_{e} with a Fermi–Dirac distribution, and we sample a Brillouin Zone with a 160 × 160 × 1 mesh. This does not take into account anharmonic effects, assuming T_{l} = 300K. Figure 7 shows a weak ΔPos(2D) ∼5 cm^{−1} in the range T_{e} = 300–3000 K. In equilibrium, T_{l} = T_{e} would induce a nonnegligible anharmonicity^{72}, which would lead to a Pos(2D) softening: ΔPos(2D)/ΔT_{eq} ≈ −0.05 cm^{−1} K^{−1}. The weak dependence ΔPos(2D)(P_{L}) (blue circles in Fig. 7) rules out a dominant anharmonicity contribution and, consequently, T_{l} = T_{e}. The minor disagreement with DFPT suggests a T_{l} slightly larger than RT, but definitely smaller than T_{eq}.
Data availability
All relevant data are available from the authors.
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Acknowledgements
We acknowledge funding from the Graphene Flagship, ERC Grant Hetero2D, EPSRC Grants EP/K01711X/1, EP/K017144/1, EP/N010345/1, EP/L016087/1 and MAECI under the ItaliaIndia collaborative project SuperTopPGR04879.
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T.S. led the research project, conceived with G.C., F.M. and A.C.F.; C.Fe., A.V., and T.S. designed and built the pulsed Raman setup; C.Fe. and A.V. performed the out of equilibrium Raman experiments, with contribution from M.M.; C.Fa., P.P., D.D.F., U.S., A.K.O., and D.Y. performed the equilibrium CW Raman experiment; L.B. and F.M. developed the modeling and carried out the numerical simulations, with contributions from A.V.; D.D.F., U.S., A.K.O., and D.Y. prepared and characterized the samples; C.Fe., A.V., L.B., G.C., F.M., A.C.F., and T.S. interpreted the data and the simulations and wrote the manuscript.
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Ferrante, C., Virga, A., Benfatto, L. et al. Raman spectroscopy of graphene under ultrafast laser excitation. Nat Commun 9, 308 (2018). https://doi.org/10.1038/s4146701702508x
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