Generation of highly-polarized high-energy brilliant γ-rays via laser-plasma interaction

Kun Xue; Zhen-Ke Dou; Feng Wan; Tong-Pu Yu; Wei-Min Wang; Jie-Ru Ren; Qian Zhao; Yong-Tao Zhao; Zhong-Feng Xu; Jian-Xing Li

doi:10.1063/5.0007734

Abstract

The generation of highly polarized high-energy brilliant γ-rays via laser–plasma interaction is investigated in the quantum radiation-reaction regime. We employ a quantum electrodynamics particle-in-cell code to describe spin-resolved electron dynamics semiclassically and photon emission and polarization quantum mechanically in the local constant field approximation. As an ultrastrong linearly polarized (LP) laser pulse irradiates a near-critical-density (NCD) plasma followed by an ultrathin planar aluminum target, the electrons in the NCD plasma are first accelerated by the driving laser to ultrarelativistic energies and then collide head-on with the laser pulse reflected by the aluminum target, emitting brilliant LP γ-rays via nonlinear Compton scattering with an average polarization of about 70% and energy up to hundreds of MeV. Such γ-rays can be produced with currently achievable laser facilities and will find various applications in high-energy physics and laboratory astrophysics.

I. INTRODUCTION

Polarized high-energy γ-rays have a wide range of significant applications, for example, in generating polarized positrons and electrons,^1,2 probing radiation mechanisms and properties of dark matter³ and black holes,⁴ exciting polarization-dependent photofission of nuclei in giant dipole resonances,⁵ meson photoproduction,⁶ and detecting vacuum birefringence in ultrastrong laser fields.^7–11 Such γ-rays are commonly produced through either bremsstrahlung^12,13 or linear Compton scattering.^14–16 However, the former cannot generate linearly polarized (LP) γ-rays, suffers from the deficiencies of large scattering angle and divergent emission in the incoherent regime,¹⁷ and is limited to a low current density of the impinging electrons and a low radiation flux owing to the risk of damage to crystalline materials in the coherent regime,^18–20 while the latter is severely restricted by the low electron–photon collision luminosity due to the low laser intensities.

Nowadays, with rapid developments in high-power laser technology, state-of-the-art laser facilities can provide laser beams with peak intensities of the order of 10²² W/cm², pulse durations of tens of femtoseconds, and energy fluctuations of ∼1%,^21–27 which have stimulated experimental investigations of quantum electrodynamics (QED) processes during laser–plasma or laser–electron beam interactions.^28,29 Such strong laser fields can be employed to directly polarize electrons^30–36 through radiative spin effects and to create polarized positrons^37,38 through the asymmetry in spin-resolved pair production probabilities. Moreover, in such strong laser fields, the Compton scattering process moves into the nonlinear regime: during laser–electron interaction, the electron radiates a high-energy γ-photon by absorbing millions of laser photons.²⁸ Highly polarized high-energy brilliant γ-rays can be generated via nonlinear Compton scattering in laser–electron beam interaction.^39–41 In ultrastrong laser fields, the radiation formation length (inversely proportional to the laser intensity) is much smaller than the laser wavelength and cannot carry the driving laser helicity. Thus, the circular polarization of the emitted γ-photons is transferred from the angular momentum (helicity) of electrons⁴¹ (by contrast, the generation of LP γ-photons does not require electron polarization⁴²). This mechanism is very different from that in the regimes of Thomson scattering^43,44 and of linear and weakly nonlinear Compton scattering,^45,46 in which the polarization of the emitted γ-photons is transferred from the driving laser field since the radiation formation length is longer than the laser wavelength. Further interaction of the generated polarized high-energy γ-photons with the laser fields could produce electron–positron pairs via the multiphoton Breit–Wheeler process,^47–50 in which case the photon polarization will have a significant effect on the pair production probability.^{17,42,47,51–54}

Recently, all-optical γ-photon sources have attracted widespread interest.^55–58 Most of the methods for generating brilliant high-energy γ-rays that have either been experimentally demonstrated or proposed theoretically are based on bremsstrahlung,^59,60 nonlinear Thomson scattering,^57,58 the synchrotron effect,^61,62 betatron oscillations,^63,64 electron wiggling,⁶⁵ and nonlinear Compton scattering.^66–69 However, in these innovative works, information about γ-photon polarization has generally been overlooked, in particular, in the regime of strongly nonlinear Compton scattering in the laser–plasma interaction, which actually plays a significant role in subsequent secondary particle generation.^{17,42,47,51–54} Therefore, the polarization process of the γ-photons emitted during laser–plasma interaction remains an open question.

In this paper, highly polarized high-energy brilliant γ-rays generated by laser–plasma interaction are studied in the quantum radiation-reaction regime with currently achievable laser intensities.^21–27 We implement electron spin and photon polarization algorithms in two-dimensional (2D) and three-dimensional (3D) particle-in-cell (PIC) EPOCH codes^70,71 in the local constant field approximation^{17,49,72–74} to describe spin-resolved electron dynamics semiclassically and photon emission and polarization quantum mechanically.^{34,37,38,41,42} We consider a commonly used experimental setup: an ultrastrong LP laser pulse irradiating a near-critical-density (NCD) hydrogen plasma followed by an ultrathin planar aluminum (Al) target (see the interaction scenario in Fig. 1). The electrons in the plasma are first accelerated by the driving laser pulse to ultrarelativistic energies and then collide head-on with the laser pulse reflected by the Al target, emitting abundant LP γ-photons via nonlinear Compton scattering with an average polarization of about 70%, energy up to hundreds of MeV, and brilliance of the order of 10²¹ photons/(s mm² mrad² 0.1%BW) for the given parameters. Compared with laser–electron beam interaction,⁴¹ this scenario is more easily accessible, since only a single laser beam is required. We also show the impact of the laser and target parameters on the γ-ray polarization and brilliance. As a conical gold (Au) target filled with an NCD hydrogen plasma is employed instead of a planar target, for the same driving laser, the collimation and energy of emitted γ-rays are improved, but the polarization rate is reduced, because of the impact of the laser-driven strong quasi-static magnetic field [see the interaction scenario in Fig. 4(a)].

Figure 1.Scenario for the generation of LP γ-rays via nonlinear Compton scattering. An ultrastrong LP laser pulse, polarized along the y axis and propagating along the x direction, irradiates an NCD hydrogen plasma followed by an ultrathin planar Al target. Laser-driven ultrarelativistic electrons in the plasma collide head-on with the laser pulse reflected by the Al target, emitting LP high-energy γ-photons, which can penetrate through the Al target and propagate forward.

II. SIMULATION METHOD

In this work, we consider laser–plasma interaction in the quantum radiation-reaction regime, which requires a large nonlinear QED parameter $χ_{e} \equiv | e | \sqrt{- {(F_{μ ν} p^{ν})}^{2}} / m^{3} ≳ 1$ .^49,75 Multiphoton Breit–Wheeler pair production is characterized by another nonlinear QED parameter $χ_{γ} \equiv | e | \sqrt{- {(F_{μ ν} k_{γ}^{ν})}^{2}} / m^{3}$ .^17,49 Here, p and k_γ are the 4-momenta of the electron and photon, respectively, e and m are the electron charge and mass, respectively, and F_μν is the field tensor. As the electron collides head-on with the laser, one can make the estimate χ_e ≈ 2a₀γ_eω₀/m, with γ_e being the electron Lorentz factor and a₀ = eE₀/mω₀ the invariant laser field parameter. Here, E₀ and ω₀ are the laser field amplitude and frequency, respectively. Relativistic units with c = ℏ = 1 are used throughout.

In our simulation method, we treat spin-resolved electron dynamics semiclassically, and photon emission and pair production quantum mechanically in the local constant field approximation,^{17,49,72–74} which is valid at a₀ ≫ 1. In each simulation step, the photon emission and polarization are both electron-spin-dependent and are calculated using Monte Carlo algorithms^34,41,42 derived in the leading-order approximation with respect to 1/γ_e via the QED operator method of Baier and Katkov.⁷⁶ The photon polarization is represented by the Stokes parameters (ξ₁, ξ₂, ξ₃), defined with respect to the axes ${\hat{e}}_{1} = \hat{a} - \hat{v} \cdot (\hat{v} \cdot \hat{a})$ and ${\hat{e}}_{2} = \hat{v} \times \hat{a}$ ,⁷⁷ with the photon emission direction $\hat{n}$ being along the electron velocity v for an ultrarelativistic electron (the emission angle ∼1/γ_e ≪ 1). Here, $\hat{v} = v / | v |$ , and the unit vector $\hat{a} = a / | a |$ is along the electron acceleration a. On detecting the mean polarization of a γ-photon beam, one must first normalize the Stokes parameters of each photon to the same observation frame ( ${\hat{o}}_{1}$ , ${\hat{o}}_{2}$ , $\hat{n}$ ), i.e., rotate the Stokes parameters of each photon from its instantaneous frame ( ${\hat{e}}_{1}$ , ${\hat{e}}_{2}$ , $\hat{n}$ ) to the same observation frame ( ${\hat{o}}_{1}$ , ${\hat{o}}_{2}$ , $\hat{n}$ ), and then calculate the average Stokes parameters of the γ-photon beam.^41,42,77 Alternatively, in quantum mechanics, the density matrix of photons in terms of the pure state e_α is given by $ρ_{α β} \equiv e_{α} e_{β}^{*}$ , and it can also be expressed as the following combination of the Stokes parameters and the Pauli matrix σ:⁷⁷

After photon emission, the electron spin state is determined by the spin-resolved emission probabilities and instantaneously collapses into one of its basis states defined with respect to the instantaneous spin quantization axis (SQA), which is chosen according to the particular observable of interest: to determine the polarization of the electron along the magnetic field in its rest frame, the SQA is chosen along the direction of the magnetic field, $n_{B} = \hat{v} \times \hat{a}$ ;^34,37 in the case when the electron beam is initially polarized with initial spin vector S_i, the observable of interest is the spin expectation value along the initial polarization, and the SQA is chosen along that direction.⁴¹ Between photon emissions, the spin precession is governed by the Thomas–Bargmann–Michel–Telegdi equation.^78–80 One can, alternatively, use the Cain code⁸¹ to obtain uniform results.

As emitted high-energy γ-photons interact further with the strong laser fields, electron–positron pairs could be produced by the multiphoton Breit–Wheeler process,^47–50 and the pair production probabilities depend on the photon polarization,^{17,47,51–54} which can be calculated using the Monte Carlo method.⁴²

III. POLARIZATION OF γ-RAYS GENERATED BY LASER–PLASMA INTERACTION

An ultrastrong LP laser pulse irradiates an NCD hydrogen plasma followed by an ultrathin planar Al target (Fig. 1). Electrons in the plasma are accelerated to ultrarelativistic energies and then collide head-on with the laser pulse reflected by the Al target, emitting abundant LP high-energy γ-photons. To maximize the reflection of the driving laser pulse, the thickness of the Al target should be greater than the laser piston depth.⁸² In fact, as the emitted γ-photons penetrate through the Al target, an appropriate thickness can also mitigate bremsstrahlung and Bethe-Heitler pair production, which usually requires a thickness of the order of millimeters.^83,84

The simulation box employed is x × y = 40λ₀ × 30sλ₀, and the corresponding cells are 1000 × 750. An LP laser pulse injected from the left boundary polarizes along the y axis and propagates along the x direction with wavelength λ₀ = 1 µm and normalized intensity $a = a_{0} \exp [- {(t - t_{0})}^{2} / τ^{2}] \exp (- y^{2} / w_{0}^{2})$ , where the focal radius w₀ = 5λ₀ and the pulse duration τ = 9T₀, with laser period T₀, corresponding to a full width at half maximum (FWHM) $τ^{'} = 2 \sqrt{\ln 2} τ \approx 15 T_{0}$ , and a time delay t₀ = τ is adopted. A solid planar Al target with electron density $n_{e}^{Al} = 702 n_{c}$ and thickness d_Al = 1 µm is placed 20 µm from the right boundary, where the plasma critical density $n_{c} = m ω_{0}^{2} / 4 π e^{2} \approx 1.1 \times 1 0^{21}$ cm⁻³. The left side of the Al target is filled with an NCD hydrogen plasma with electron density n_e = 5n_c and thickness d_p = 10 µm. The numbers of macroparticles in each cell are 100 for electrons and 20 for H⁺ and for Al¹³⁺ (fully ionized).

Distributions of the density and linear polarization of emitted γ-rays are illustrated in Figs. 2(a) and 2(b), respectively, with the laser peak intensity I₀ ≈ 1.38 × 10²² W/cm² (a₀ = 100).^21–27 High-energy γ-photons with ɛ_γ ≳ 100 MeV are mainly emitted forward, and two density peaks arise near θ ≈ ±21°, since the γ-photons are assumed to be emitted along the electron momentum, and the electron propagation angle $θ_{e} \propto p_{e}^{⊥} / p_{e}^{∥} \propto a_{R} / γ_{e}$ (the electrons interact with the reflected laser pulse with invariant intensity parameter a_R and $a_{R} \propto \bar{a} \propto a_{0}$ with an average laser intensity $\bar{a}$ ), as shown in Fig. 2(a), which is in excellent agreement with other simulations.⁸⁵ The linear polarization of γ-photons is characterized by the Stokes parameters ξ₁ and ξ₃.^41,42,77 As we employ the basis vector of the observation frame ${\hat{o}}_{1}$ in the polarization y–x plane of the driving laser, and because γ-photons are emitted mainly in the polarization plane, ξ₁ is negligible and ξ₃ is as given in Fig. 2(b): at a given polar angle, the linear polarization ξ₃ for higher-energy γ-photons is larger. The average linear polarization $\bar{ξ_{3}} \approx 0.68$ , and the partial ξ₃ can reach up to 0.73. In nonlinear Compton scattering, the circular polarization defined by ξ₂ requires initially longitudinally spin-polarized electrons⁴¹ and is negligible here.

Figure 2.(a) Angle-resolved density of emitted γ-rays log₁₀[d²N_γ/dε_γdθ (MeV⁻¹ mrad⁻¹)] vs γ-photon energy ε_γ and polar angle θ. (b) Linear polarization ξ₃ of emitted γ-photons vs ε_γ and θ. In (a) and (b), a₀ = 100. [(c) and (d)] dN_γ/dθ and average linear polarization $\bar{ξ_{3}}$ , respectively, vs θ, calculated by summing over ε_γ in (a) and (b), respectively. (e) dN_γ/dε_γ calculated by summing over θ in (a) vs ε_γ. The red, blue, and black curves in (c)–(e) are for the cases a₀ = 50, 100, and 150, respectively. Only γ-photons with ε_γ ≥ 1 MeV are counted. Other laser and target parameters are given in the text. (f) ξ₃ vs χ_e and ε_γ/ε_e.

To make the results in Figs. 2(a) and 2(b) clearer, we sum over ε_γ to obtain the angle-resolved number and linear polarization of emitted γ-photons shown by the blue curves in Figs. 2(c) and 2(d), and we sum over θ in Fig. 2(a) to obtain the energy density shown by the blue curve in Fig. 2(e). The impact of the driving laser intensity a₀ can be seen from Figs. 2(c)–2(e). In ultrastrong laser fields, in one laser period T₀, the number of formation lengths (in which a photon could be emitted) is proportional to the invariant laser intensity, and the photon emission probability in each formation length is proportional to the fine structure constant α,^28,49 and thus the number of emitted γ-photons N_γ ∝ N′_eαa_Rτ/T₀ ∝ a₀, where N′_e is the number of parental electrons and is proportional to the target density n_e and laser focal spot size w₀: $N'_{e} \propto n_{e} π w_{0}^{2}$ . Since a₀ ≫ 1, nearly all atoms within the focal spot can be fully ionized. As a₀ increases from 50 to 150, N_γ also continuously increases, the reflected laser intensity a_R and the electron energy γ_e are both increased, and consequently the peak angle θ ∝ a_R/γ_e increases slightly ^{[Fig. 2(c)]}. ε_γ ∝ ε_eχ_e, where the electron energy ε_e ∼ a₀ (i.e., a stronger driving laser can accelerate electrons to higher energies) and χ_e ∝ a_Rγ_e ∝ a₀ε_e, and thus, as a₀ increases, ε_γ increases as well ^{[Fig. 2(e)]}.

We underline that $\bar{ξ_{3}}$ decreases monotonically with increasing a₀, as shown in Fig. 2(d). The physical reason for this is as follows. The Stokes parameter ξ₃ depends sensitively on the parameters χ_e and ε_γ/ε_e (see the analytical expression for ξ₃ in Refs. 41 and 42), as demonstrated in Fig. 2(f). As χ_e increases in the region under consideration, ξ₃ decreases continuously; as ε_γ/ε_e increases, ξ₃ first increases slightly and then gradually decreases to 0. Consequently, as χ_e ∝ a₀ increases, $\bar{ξ_{3}}$ decreases in Fig. 2(d); i.e., lower-intensity driving laser pulses can generate higher-polarization (but lower-brilliance) γ-photons.

In the case a₀ = 100, the radius and duration of the emitted γ-ray beam are w_γ ≈ w₀ and τ_γ ≈ τ′, respectively. The angular divergences (FWHM) are approximately 1.74 × 1.74 rad,² 1.47 × 1.47 rad², 1.39 × 1.39 rad², and 1.36 × 1.36 rad² for ε_γ ≥ 1 MeV, 10 MeV, 100 MeV, and 200 MeV, respectively. The corresponding brilliances are 0.67 × 10²¹, 2 × 10¹⁹, 7 × 10¹⁸, and 2 × 10¹⁷ photons/(s mm² mrad² 0.1% BW) for ε_γ = 1 MeV, 10 MeV, 100 MeV, and 200 MeV, respectively. It is obvious that the brilliance ∝ N_r ∝ a₀. For the given parameters, the multiphoton Breit–Wheeler pair production probabilities during the interactions of γ-photons with the laser fields are rather low, since χ_γ ≪ 1.

Note that for the given parameters, as the γ-photons propagate through the plasma, the photon polarization flips (depolarization) in the background fields (vacuum polarization), and the collisions with the plasma particles (involving Compton scattering and Bethe–Heitler pair production) are estimated to be negligible.^17,86

With the aim of determining experimental feasibility, the impact of laser and plasma parameters on the density and polarization degree of emitted γ-photons is shown in Fig. 3. For instance, as the plasma density n_e or the driving laser pulse duration τ increases, N_γ ∝ N′_eαa_Rτ/T₀ ∝ n_eτ also increases ^{[Figs. 3(a) and 3(c)]}. However, for an NCD plasma, $\bar{ξ_{3}}$ changes only slightly with variations in n_e and τ ^{[Figs. 3(b) and 3(d)]}. Note that as n_e increases from 2n_c to 10n_c, the opacity of the plasma increases as well, and laser propagation becomes more and more unstable, which results in an asymmetric angular distribution of emitted γ-rays,⁶¹ as shown in Fig. 3(a). Meanwhile, in the plasma, as a result of the driving laser pulse pushing the electrons in the low-density region forward, a high-density region can be created, where unstable laser propagation further induces asymmetric γ-photon emission. This effect becomes more severe as the laser pulse duration τ increases, as shown in Fig. 3(c). As already shown in Fig. 2, the peak angle of the γ-ray spectrum θ ∝ a_R/γ_e, and a_R and γ_e here both depend on the laser and plasma parameters (e.g., a₀, τ, and n_e). It is difficult to obtain analytical relations between θ and these parameters; however, our simulation results in Figs. 2(c), 3(a), and 3(c) indicate that the peak angle θ changes only slightly when these parameters vary, which is advantageous for experimental observations.

Figure 3.(a) and (b) dN_γ/dθ and $\bar{ξ_{3}}$ , respectively, vs θ for the cases n_e = 2n_c (red), 5n_c (blue), and 10n_c (black). [(c) and (d)] dN_γ/dθ and $\bar{ξ_{3}}$ , respectively, vs θ for the cases τ = 6T₀ (red), 9T₀ (blue), and 12T₀ (black). a₀ = 100, and other laser and target parameters are the same as those in Fig. 2.

Furthermore, instead of the planar target in Figs. 1 and 2, we now consider a conical Au target filled with an NCD hydrogen plasma ^{[Fig. 4(a)]}. The front and rear surfaces of the cone are open. As an ultraintense LP laser pulse irradiates the NCD hydrogen plasma inside the Au cone, almost all the bulk electrons are pushed forward and excite a strong quasi-static magnetic field B_p.^61,66 The maximum intensity of the magnetic field can be estimated as max(B_p) ≃ 4π|e|β_en_eR, which is of the same order of magnitude as the magnetic field of the driving laser.⁶¹ Here, R and β_e denote the radius of the cone and the electron velocity scaled by the light speed in vacuum. For the plane-wave case, the transverse electric field E_⊥ can almost cancel v · B with the magnetic field B, and consequently χ_e is rather small and the photon emission is very weak. However, in the case of a conical target with a strong quasi-static magnetic field B_p, χ_e ∝ B_p greatly increases, and thus subsequent γ-photon emission is significantly enhanced. Furthermore, since most photons are emitted at the edge of the cone, where the transverse velocities of electrons are close to 0, the angular spread of the γ-ray beam is narrowed.⁶¹

Figure 4.(a) Scenario for the generation of LP γ-rays by an ultrastrong LP laser pulse, polarizing along the y axis and propagating along the x direction, interacting with a conical Au target filled with an NCD hydrogen plasma. The red and blue areas inside the cone indicate the quasi-static magnetic fields along the +z and −z directions, respectively, caused by the driving laser pulse. (b) Plot of log₁₀(dN_γ/dΩ) with respect to the polar angle θ and the azimuthal angle ϕ, where the solid angle dΩ = sin θ dθ dϕ. (c) Plot of the linear polarization P_LP with respect to θ and ϕ, where $P_{LP} = \sqrt{ξ_{1}^{2} + ξ_{3}^{2}}$ . (d) Total energy spectra of emitted γ-photons dN_γ/dε_γ vs ε_γ. (e) dN_γ/dε_γ (black) and P_LP (red) vs ε_γ at angles (θ, ϕ) = (10° ± 0.3°, 180° ± 0.3°), where P_LP is the average linear polarization of γ-photons with energies ≥ ε_γ. The results in (b)–(e) are calculated with ε_γ ≥ 1 MeV at t = 110T₀, when the interaction has finished. The laser parameters are the same as those in Figs. 2(a) and 2(b), and the target parameters are given in the text.

In Fig. 4, the simulation box is x × y × z = 110λ₀ × 40λ₀ × 40λ₀, and the corresponding cells are 2200 × 400 × 400. A solid conical Au target with electron density $n_{e}^{Au} = 100 n_{c}$ and thickness d_Au = 1 µm is located in the region from 5λ₀ to 75λ₀ on the x axis. The left and right opening radii of the cone are R₁ = 7λ₀ and R₂ = 3λ₀, respectively. The Au cone is filled with an NCD hydrogen plasma of density n_e = 5n_c. The numbers of macroparticles in each cell are 100 for electrons and 20 for H⁺ and for Au²⁺ (partially ionized). The angle-resolved density of emitted γ-photons is illustrated in Fig. 4(b). Compared with the case of a planar target employing the same driving laser, the collimation and cutoff energy of the emitted γ-photons in the case of a conical target are much better, since the conical target has a greater longitudinal size and generates strong confinement effects derived from the transverse magnetic field. For instance, most γ-photons concentrate in the angular region θ ≲ 20°, and the cutoff energy exceeds 450 MeV, as shown in Fig. 4(d); see for comparison the corresponding values for a planar target in Fig. 2. The azimuthal magnetic field generated in the conical target results in a transverse (azimuthal) spread of the electron dynamics and further γ-photon emission. This reduces the average polarization rate of the γ-rays, as shown in Fig. 4(c), because γ-photons polarized along different azimuthal angles cancel each other in the average polarization. For instance, the energy density and polarization of the γ-photons at a specific solid angle (θ, ϕ) = (10° ± 0.3°, 180° ± 0.3°) near the γ-ray beam center are illustrated in Fig. 4(e). The linear polarization and the energy density are respectively proportional and inversely proportional to the γ-photon energy. For ε_γ ≥ 20 MeV, 40 MeV, and 60 MeV, the linear polarizations P_LP ≈ 0.28, 0.41, and 0.52, respectively, and the brilliances are approximately 0.86 × 10²⁰, 0.43 × 10²⁰, and 0.30 × 10²⁰ photons/(s mm² mrad² 0.1% BW), respectively.

IV. CONCLUSION

Generation of highly polarized high-energy brilliant γ-rays has been studied through investigating the interaction of a single-shot ultraintense LP laser pulse with a solid target filled with an NCD plasma in the quantum radiation-reaction regime. The results show that with currently achievable laser intensities, the emitted γ-rays could reach an average linear polarization of about 70%, an energy up to hundreds of MeV, and a brilliance of the order of 10²¹ photons/(s mm² mrad² 0.1% BW), and could have applications in high-energy physics and laboratory astrophysics.

AUTHORS’ CONTRIBUTIONS

K.X. and Z.-K.D. contributed equally to this work.

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