Hyperfine Structure of the $B^3\Pi_1$ State and Predictions of Optical Cycling Behavior in the $X\rightarrow B$ transition of TlF

TL;DR

This paper measures the hyperfine structure of TlF molecules and predicts their optical cycling behavior, providing data crucial for laser cooling and precision measurements involving TlF.

Contribution

It provides new measurements of hyperfine and rotational constants and predicts optical cycling efficiency in TlF, advancing molecular laser cooling research.

Findings

Measured hyperfine and rotational constants for TlF B state.

Updated vibrational branching fractions for TlF.

Predicted photon scattering limits for optical cycling schemes.

Abstract

The rotational and hyperfine spectrum of the $X^1\Sigma^+ \rightarrow B^3\Pi_1$ transition in TlF molecules was measured using laser-induced fluorescence from both a thermal and a cryogenic molecular beam. Rotational and hyperfine constants for the $B$ state are obtained. The large magnetic hyperfine interaction of the Tl nuclear spin leads to significant mixing of the lowest $B$ state rotational levels. Updated, more precise measurements of the $B\rightarrow X$ vibrational branching fractions are also presented. The combined rovibrational branching fractions allow for the prediction of the number of photons that can be scattered in a given TlF optical cycling scheme.

Tables

Table 1. Table 1: Observed splittings of e 𝑒 e -parity levels (MHz). Regular (italic) type indicates splittings measured in the thermal (cryogenic) beam.

$\tilde{\text{J}}$	R	a	b	c	d	e	f	g		P	a	b	c	d	e	f	g
21	20	37	354	42
…
16	15	50	303	60
…
11	10	66	274	83
…
9	8	91	272	99
8	7	88	270	110	2203	80	224	113
7	6	93	269	128	2157	93	224	128
6	5	105	297	153	2108	113	231	150
5	4	121	304	189	2050	125	247	185
4	3	147	343	255	1909	147	266	249		5	142	379	250	1813	151	254	247
3	2	178	391	341	1735	179	302	335		4	175	388	341	1646	188	309	332
2	1	227	463	580	1378	229	326	579		3	256	475	562	1340	225	315	560
1	0	548	13524	304				289		2	545	13523	305	-11711	582	13431	309

Table 2. Table 2: Observed splittings of f 𝑓 f -parity levels (MHz). Regular (italic) type indicates splittings measured in the thermal (cryogenic) beam. Bold indicates splittings of tentative line assignments based on comparisons of overlapping lines present in both beams.

$\tilde{\text{J}}$	Q	a	b	c	d	e	f	g	O	h	i	j	S	k	l	m
70	70						177
69	69						176
68	68		179				171
67	67						173
66	66		175				168
65	65		176				168
64	64		175				167
63	63		184				166
62	62		177				165
61	61		170				164
60	60		168
59	59		157
…
51	51	17	139	17
…
34	34	27	111	26
…
29	29	26	114	26
…
27	27	25	113	32
…
25	25	30	108	29
…
23	23	37	110	42
…
20	20	36	110	45
…
17	17	46	115	50
16	16	48	115	54
…
14	14	56	121	64
13	13	59	133	70
12	12	63	131	76
11	11	68	140	80
10	10	74	157	78
9	9	75	155	100
8	8	90	164	111
7	7	93	194	116
6	6	107	221	150									4			154
5	5	129	255	179									3	180	2430	184
4	4	152	302	245									2	244	2355	235
3	3	182	372	344									1	367	2229	347
2	2	233	472	568									0	561	1948	578
1	1	571	13522	316	-11361	553			3	312	2269	309

Table 3. Table 3: Fit rotational and HF parameters for the B 3 ​ Π 1 superscript 𝐵 3 subscript Π 1 B^{3}\Pi_{1} state of TlF. Numbers in parentheses are the 1 σ 𝜎 \sigma confidence intervals. All values are in MHz. The root mean square deviation between our fitted and measured splittings is < 8 absent 8 <8  MHz in all cases. The residual is < 20 absent 20 <20  MHz for each fitted point, and tends to decrease with increasing J ~ ~ 𝐽 \tilde{J} .

	This Work			Ref. Tiemann (1988)
	$e$-parity		$f$-parity	$e$-parity		$f$-parity
Parameter	203TlF	205TlF	205TlF	203TlF	205TlF	203TlF	205TlF
$B_0$	6694.42(19)	6689.09(16)	6686.667(26)	6694.5(11)	6688.9(11)	6694.3(11)	6688.7(11)
$D_0$			0.010869(27)	0.01089(14)	0.01088(14)	0.01098(14)	0.01096(14)
$H_0\cdot {10}^8$			-8.1(6)	-9.0(6)	-9.0(6)	-9.0(6)	-9.0(6)
$h_1$(Tl)	28516(5)	28793(5)	28802(4)
$h_1$(F)	864(1)	859(1)	871(1)
$c_I$(Tl)	-13.9(1)	-13.5(1)	-2.2(1)

Table 4. Table 4: Eigenstates and their rotational branching ratios r J ~ ​ J g subscript 𝑟 ~ 𝐽 subscript 𝐽 𝑔 r_{\tilde{J}J_{g}} for | B ​ ( v e = 0 ) , F = 0 − 3 , P = f ⟩ ket formulae-sequence 𝐵 subscript 𝑣 𝑒 0 𝐹 0 3 𝑃 𝑓 \ket{B(v_{e}=0),F=0-3,P=f} , calculated from the best-fit spectroscopic constants in Tables 2 and 3 for 205 TlF.

Nominal State Label	Calculated Admixture $|J,F_1,F⟩$				$r_{\tilde{J}\tilde{J}-2}$	$r_{\tilde{J}\tilde{J}}$	$r_{\tilde{J}\tilde{J}+2}$
$|\tilde{J}=1,F_1=\frac{1}{2},F=0⟩$	$|1,\frac{1}{2},0⟩$					1
$|\tilde{J}=1,F_1=\frac{1}{2},F=1⟩$	$+0.9996|1,\frac{1}{2},1⟩$	$+0.0203|1,\frac{3}{2},1⟩$	$-0.0180|2,\frac{3}{2},1⟩$			0.9998	0.0002
$|\tilde{J}=1,F_1=\frac{3}{2},F=1⟩$	$+0.0267|1,\frac{1}{2},1⟩$	$-0.8518|1,\frac{3}{2},1⟩$	$+0.5232|2,\frac{3}{2},1⟩$			0.836	0.164
$|\tilde{J}=2,F_1=\frac{3}{2},F=1⟩$	$+0.0048|1,\frac{1}{2},1⟩$	$+0.5235|1,\frac{3}{2},1⟩$	$+0.8520|2,\frac{3}{2},1⟩$		0.091	0.909
$|\tilde{J}=1,F_1=\frac{3}{2},F=2⟩$	$+0.8482|1,\frac{3}{2},2⟩$	$-0.5294|2,\frac{3}{2},2⟩$	$-0.0138|2,\frac{5}{2},2⟩$	$+0.0064|3,\frac{5}{2},2⟩$		0.832	0.168
$|\tilde{J}=2,F_1=\frac{5}{2},F=2⟩$	$-0.0104|1,\frac{3}{2},2⟩$	$+0.0120|2,\frac{3}{2},2⟩$	$-0.9353|2,\frac{5}{2},2⟩$	$+0.3535|3,\frac{5}{2},2⟩$	4$\cdot {10}^{-5}$	0.929	0.071
$|\tilde{J}=2,F_1=\frac{3}{2},F=2⟩$	$+0.5295|1,\frac{3}{2},2⟩$	$+0.8482|2,\frac{3}{2},2⟩$	$+0.0011|2,\frac{5}{2},2⟩$	$-0.0103|3,\frac{5}{2},2⟩$	0.094	0.906	6$\cdot {10}^{-5}$
$|\tilde{J}=3,F_1=\frac{5}{2},F=2⟩$	$+0.0040|1,\frac{3}{2},2⟩$	$+0.0085|2,\frac{3}{2},2⟩$	$+0.3536|2,\frac{5}{2},2⟩$	$+0.9353|3,\frac{5}{2},2⟩$	0.050	0.950
$|\tilde{J}=2,F_1=\frac{5}{2},F=3⟩$	$+0.9341|2,\frac{5}{2},3⟩$	$-0.3568|3,\frac{5}{2},3⟩$	$-0.0100|3,\frac{7}{2},3⟩$	$+0.0032|4,\frac{7}{2},3⟩$		0.927	0.073
$|\tilde{J}=3,F_1=\frac{7}{2},F=3⟩$	$-0.0084|2,\frac{5}{2},3⟩$	$+0.0074|3,\frac{5}{2},3⟩$	$-0.9638|3,\frac{7}{2},3⟩$	$+0.2665|4,\frac{7}{2},3⟩$	3$\cdot {10}^{-5}$	0.961	0.039
$|\tilde{J}=3,F_1=\frac{5}{2},F=3⟩$	$+0.3568|2,\frac{5}{2},3⟩$	$+0.9341|3,\frac{5}{2},3⟩$	$+0.0017|3,\frac{7}{2},3⟩$	$-0.0084|4,\frac{7}{2},3⟩$	0.051	0.949	4$\cdot {10}^{-5}$
$|\tilde{J}=4,F_1=\frac{7}{2},F=3⟩$	$+0.0023|2,\frac{5}{2},3⟩$	$+0.0073|3,\frac{5}{2},3⟩$	$+0.2666|3,\frac{7}{2},3⟩$	$+0.9638|4,\frac{7}{2},3⟩$	0.031	0.970

Table 5. Table 5: Predicted and measured values of the branching fractions b 0 ​ v g subscript 𝑏 0 subscript 𝑣 𝑔 b_{0v_{g}} in TlF. The calculations were done separately using the Morse potential and the RKR potential as models for the internuclear potential.

$B\left(0\right)\to X\left(v_g\right)$	Morse	RKR	Measured
$v_g$=0 (271.7 nm)	0.9892(3)	0.989(2)	0.989(2)
$v_g$=1 (275.3 nm)	0.0003(2)	0.0005(3)	$15\left(4\right)\cdot {10}^{-5}$
$v_g$=2 (278.8 nm)	0.0104(2)	0.010(2)	$0.011\left(2\right)$
$v_g$=3 (282.5 nm)	0.00000(1)	0.0003	$3\left(2\right)\cdot {10}^{-5}$
$v_g$=4 (286.2 nm)	0.00013(1)	0.0002	$13\left(3\right)\cdot {10}^{-5}$
$v_g$=5 (290.0 nm)	0.00000(1)	0.0003	$1\left(1\right)\cdot {10}^{-5}$
$v_g$=6 (293.8 nm)	$-$	0.0002	$1\left(2\right)\cdot {10}^{-5}$

Table 6. Table 6: Part numbers (Andover), transmissions, and bandwidths of all the bandpass filters used in this experiment. Uncertainty in the transmissions of all filters except the neutral density filter is 1.5%. Uncertainty in transmission of the neutral density filter at the given wavelength is 1%. Uncertainties were chosen to reflect the range of possible transmissions that could result with incidences between ± plus-or-minus \pm 2 ∘ off normal.

$B\left(0\right)\to X\left(v_g\right)$	Transition Wavelength $\lambda$	Part Number	Transmission at $\lambda$	Peak Wavelength	FWHM
$v_g=0$	271.7 nm	R214-01	17.9%	271.8 nm	10.3 nm
		R214-02	17.0%	271.3 nm	10.1 nm
		neutral density	13.5%	$-$	$-$
$v_g=1$	275.3 nm	T144-02	14.1%	275.4 nm	1.3 nm
		R088-01	13.8%	275.3 nm	1.3 nm
$v_g=2$	278.8 nm	S078-02	9.9%	279.3 nm	2 nm
$v_g=3$	282.5 nm	T236-04	12.9%	282.7 nm	1.4 nm
		T236-05	12.4%	282.8 nm	1.5 nm
$v_g=4$	286.2 nm	T338-01	11.4%	286.5 nm	1.6 nm
		T335-10	11%	290.9 nm	10.7 nm
$v_g=5$	290.0 nm	T338-02	11%	290.3 nm	1.6 nm
		T335-10	17.8%	290.9 nm	10.7 nm
$v_g=6$	293.8 nm	T335-14	11.7%	297.9 nm	10.8 nm
		T335-15	10.0%	298.7 nm	11.6 nm

Equations

$${\mathbf{F}}_1$$

$$\mathbf{F}$$

$$|c⟩=|J,\Omega ,I_1,F_1,I_2,F,m_F,P⟩.$$

$$H_{\mathrm{mhf}}=a\mathbf{I}\cdot \mathbf{L}+b\mathbf{I}\cdot \mathbf{S}+c{I}_z{S}_z,$$

$$H_{\mathrm{mhf}}^{\mathrm{eff}}=$$

$$=$$

$$\text{matrixelement}J,\Omega ,F_1,F,m{H}_{\mathrm{mhf}}(\mathrm{Tl})J^{\prime },{\Omega }^{\prime },F_1^{\prime },F,m$$

$$=h_1(\text{Tl})(-1{)}^{J+J^{\prime }+F_1+I_1-\Omega }{\delta }_{F_1,F_1^{\prime }}$$

$$\times [(2J+1)(2J^{\prime }+1)I_1(I_1+1)(2I_1+1){]}^{1/2}$$

$$\displaystyle\quad\times\begin{Bmatrix}I_{1}&J^{\prime}&F_{1}\\ J&I_{1}&1\end{Bmatrix}\matrixquantity(J\hfil&1&J^{\prime}\\ -\Omega&0&\Omega^{\prime}).$$

$$\text{matrixelement}J,\Omega ,F_1,F,m{H}_{\mathrm{mhf}}(\mathrm{F})J^{\prime },{\Omega }^{\prime },F_1^{\prime },F,m$$

$$=h_1(\text{F})(-1{)}^{2F_1^{\prime }+F+2J-\Omega +1+I_1+I_2}$$

$$\displaystyle\quad\times\begin{Bmatrix}I_{2}&F_{1}^{\prime}&F\\ F_{1}&I_{2}&1\end{Bmatrix}\begin{Bmatrix}J^{\prime}&F_{1}^{\prime}&I_{1}\\ F_{1}&J&1\end{Bmatrix}\matrixquantity(J\hfil&1&J^{\prime}\\ -\Omega&0&\Omega^{\prime})$$

$$\displaystyle\quad\times\bigl{[}(2F_{1}+1)(2F_{1}^{\prime}+1)(2J+1)(2J^{\prime}+1)$$

$$\displaystyle\quad\quad\times I_{2}(I_{2}+1)(2I_{2}+1)\bigr{]}^{1/2}.$$

$$\text{matrixelement}J,\Omega ,F_1,F,m{H}_{\mathrm{nsr}}(\mathrm{Tl})J^{\prime },{\Omega }^{\prime },F_1^{\prime },F,m$$

$$=c_I(\text{Tl})(-1{)}^{J+F_1+I_1}{\delta }_{F_1,F_1^{\prime }}{\delta }_{J,J^{\prime }}\left\{\begin{array}{ccc}{I}_1& J& F_1\\ J& I_1& 1\end{array}\right\}$$

$$\times [(J(J+1)(2J+1)I_1(I_1+1)(2I_1)+1){]}^{1/2}.$$

$$\text{matrixelement}J,\Omega ,F_1,F,m{H}_{\mathrm{nsr}}(\mathrm{F})J^{\prime },{\Omega }^{\prime },F_1^{\prime },F,m$$

$$=c_I(\text{F})(-1{)}^{2F_1^{\prime }+F+J+I_1+I_2+1}{\delta }_{J,J^{\prime }}$$

$$\times \left\{\begin{array}{ccc}{I}_2& F_1^{\prime }& F\\ F_1& I_2& 1\end{array}\right\}\left\{\begin{array}{ccc}{J}^{\prime }& F_1^{\prime }& I_1\\ F_1& J& 1\end{array}\right\}$$

$$\displaystyle\quad\times\bigl{[}(2F_{1}+1)(2F_{1}^{\prime}+1)J(J+1)(2J+1)$$

$$\displaystyle\quad\quad\times I_{2}(I_{2}+1)(2I_{2}+1)\bigr{]}^{1/2}.$$

$$H_{\mathrm{rot}}=B_0{\mathbf{J}}^2-D_0{\mathbf{J}}^4+H_0{\mathbf{J}}^6\dots ,$$

$$\text{matrixelement}J,\Omega ,F_1,F,m{H}_{\mathrm{rot}}(\mathrm{F})J,\Omega ,F_1,F,m$$

$$=B_{0}J(J+1)-D_0(J(J+1){)}^2+H_0(J(J+1){)}^3\dots .$$

$$|\tilde{J},F_1=\tilde{J}+\frac{1}{2}⟩={\alpha }_{F_1}|J=\tilde{J}⟩+{\beta }_{F_1}|J=\tilde{J}+1⟩,$$

$$|\tilde{J}+1,F_1=\tilde{J}+\frac{1}{2}⟩={\beta }_{F_1}|J=\tilde{J}⟩-{\alpha }_{F_1}|J=\tilde{J}+1⟩.$$

$$H_{2\times 2}=\matrixquantity(\frac{4F_{1}^{2}-1}{4}B+h&\frac{\sqrt{4F_{1}^{2}+4F_{1}-3}}{2(2F_{1}+1)}h\\ \frac{\sqrt{4F_{1}^{2}+4F_{1}-3}}{2(2F_{1}+1)}h&\frac{(2F_{1}+3)(2F_{1}+1)}{4}B-h).$$

$$E_{F_1}^{\pm }=B(F_1+\frac{1}{2}{)}^2\pm \frac{1}{2}\sqrt{B^2(2F_1+1{)}^2-4Bh+h^2}.$$

$$T_q^k(\mathbf{A})$$

$$=\sum _p(-1{)}^{p-q}{\mathcal{D}}_{-p-q}^{(k)}(\omega {)}^{\ast }{T}_p^k(\mathbf{A}).$$

$$\text{matrixelement}{j}_1,j_2,j_{12},m{T}^k({\mathbf{A}}_1)\cdot T^k({\mathbf{A}}_2)j_1^{\prime },j_2^{\prime },j_{12}^{\prime },m^{\prime }$$

$$=(-1{)}^{j_1^{\prime }+j_{12}+j_2}{\delta }_{j_{12},j_{12}^{\prime }}{\delta }_{m,m^{\prime }}\left\{\begin{array}{ccc}{j}_2^{\prime }& j_1^{\prime }& j_{12}\\ j_1& j_2& k\end{array}\right\}$$

$$\times \text{matrixelement}{j}_1|T^k({\mathbf{A}}_1)|j_1^{\prime }\text{matrixelement}{j}_2|T^k({\mathbf{A}}_2)|j_2^{\prime }.$$

$$\text{matrixelement}{j}_1,j_2,j_{12}|T^k({\mathbf{A}}_1)|j_1^{\prime },j_2^{\prime },j_{12}^{\prime }$$

$$=(-1{)}^{j_{12}^{\prime }+j_1+k+j_2}{\delta }_{j_2,j_2^{\prime }}\sqrt{(2j_{12}+1)(2j_{12}^{\prime }+1)}$$

$$\times \left\{\begin{array}{ccc}{j}_1^{\prime }& j_{12}^{\prime }& j_2\\ j_{12}& j_1& k\end{array}\right\}\text{matrixelement}{j}_1|T^k({\mathbf{A}}_1)|j_1^{\prime }.$$

$$\text{matrixelement}{j}_1|T^1({\mathbf{A}}_1)|j_1^{\prime }={\delta }_{j_1,j_1^{\prime }}\sqrt{j_1(j_1+1)(2j_1+1)}.$$

$$\text{matrixelement}J,\Omega |{\mathcal{D}}_{\cdot ,q}^{(k)}(\omega {)}^{\ast }|J^{\prime },{\Omega }^{\prime }$$

$$\displaystyle\quad=(-1)^{J-\Omega}\sqrt{(2J+1)(2J^{\prime}+1)}\matrixquantity(J\hfil&k&J^{\prime}\\ -\Omega&q&\Omega^{\prime}).$$

$$\text{matrixelement}J,\Omega ,F_1,F,m{T}_{q=0}^1(\mathbf{I})J^{\prime },{\Omega }^{\prime },F_1^{\prime },F,m\underset{(5.144)}{=}$$

Full text

I Introduction

The $X^{1}\Sigma^{+}$ state of thallium monofluoride (TlF) has long been a platform for precision measurements of parity- and time-reversal symmetry violation Hinds and Sandars (1980); Wilkening et al. (1984); Cho et al. (1991), with high potential for discovery of new physics Khriplovich and Lamoreaux (1997). In particular, the high mass of Tl coupled with the high polarizability of the molecule make this system ideal for measuring the Schiff moment of the Tl nucleus Schiff (1963).

Optical cycling Shuman et al. (2009) is a potentially powerful tool for enhancing the signal in a symmetry violation measurement. Unit-efficiency detection of the internal state is possible when the number of photons scattered per molecule exceeds the reciprocal of the total light collection efficiency (geometric and detector efficiencies). Optical cycling also allows for the application of large optical forces, which can be useful for precision measurements. Transverse cooling can decrease beam divergence (and increase downstream flux) Shuman et al. (2010), and longitudinal slowing and trapping can increase effective interaction times Barry et al. (2014), leading to improved energy resolution.

Optical cycling requires the ability to optically couple a subspace of ground and excited states, such that the excited states rarely decay to uncoupled ground states. Recently, we proposed the TlF $X^{1}\Sigma^{+}(v_{g}\!=\!0)$ $\rightarrow$ $B^{3}\Pi_{1}(v_{e}\!=\!0)$ transition (where $v_{g}$ ($v_{e}$) is the ground (excited) state vibrational quantum number) as a candidate for optical cycling and laser cooling Hunter et al. (2012), as this transition has a sufficiently short excited state lifetime ($\tau$ $=$ $99$ ns) and favorable vibrational branching fractions to form a quasi-closed optical cycle. In TlF, the $B^{3}\Pi_{1}$ state was expected to have resolved – and potentially very large – hyperfine (HF) structure Hunter et al. (2012). The HF interaction in the excited state can lead to mixing of rotational states with different quantum numbers $J$; this in turn can break the usual rotational selection rules, leading to branching to additional ground rotational levels which must be coupled to the optical cycle. Hence, it is crucial to characterize the rotational and HF structure of the excited state to understand and control rotational branching.

High-resolution microwave spectroscopy Hoeft et al. (1970) has provided a detailed and precise understanding of the $X$ state HF and rotational energies. Low-resolution spectroscopy with a pulsed UV laser by Wolf and Tiemann Wolf and Tiemann (1987) allowed for determination of rovibrational energies of the $B^{3}\Pi_{1}$ state. In this paper, we present high-resolution laser spectroscopy of the $X^{1}\Sigma^{+}(v_{g}=0)$ $\rightarrow$ $B^{3}\Pi_{1}(v_{e}\!=\!0)$ transition. We clearly identify rotational lines associated with $B^{3}\Pi_{1}(v_{e}=0)$ states from $J=1$ – 70. The $B$ state HF structure is fully resolved for $J\lesssim 51$ and is fit to a standard Hamiltonian to determine the parameters describing the HF interaction. These data allow a full characterization of the HF structure in the $B$ state, including effects of HF-induced rotational state mixing. In addition, we present improved vibrational branching measurements from $v_{e}=0$. Together, these data are used to quantitatively predict the number of photons that may be scattered on the $X^{1}\Sigma^{+}$ $\rightarrow$ $B^{3}\Pi_{1}$ transition of TlF for various cycling schemes.

II Experimental Details

To allow access to and identification of a large range of rotational states, we make observations using both a thermal beam source and a cryogenic buffer gas beam source. The thermal oven source is the same as in Ref. (Hunter et al., 2012). It is described in brief here. The measurements are made in a stainless steel molecular beam apparatus maintained at a pressure of about $10^{-6}$ Torr. The beam itself is created by heating a stainless steel oven containing TlF to temperatures of 688-733 K. The beam emerges from the oven through four ceramic tubes. The tubes precollimate the beam, which is then further collimated by an aperture located about 30 cm from the oven.

The cryogenic buffer gas beam source is nearly identical to that of Barry et al. (2011). A solid target is made by melting TlF powder in a copper crucible under vacuum. The crucible is mounted inside a copper cell and held at roughly 4 K by a pulse tube refrigerator. TlF molecules are produced by laser ablation of the target with 10 ns, 25 mJ pulses of 1064 nm light, and are extracted from the cell by a flow of typically 5 sccm of cryogenic helium buffer gas. The molecular beam then propagates through a region held at roughly $10^{-7}$ Torr.

The $X(0)$ $\rightarrow$ $B(0)$ transition occurs at wavelength 271.7 nm. A CW, single-frequency, tunable 1087 nm fiber laser is frequency doubled twice to produce roughly 20 mW of CW 271.7 nm light, using commercial resonant bow-tie cavities. The fiber laser frequency is locked by monitoring its transmission through a scanning Fabry-Perot cavity; the Fabry-Perot cavity length is in turn stabilized by simultaneously monitoring the transmission of a frequency-stabilized helium-neon laser.

In both molecular beam setups, the 271.7 nm laser light is directed to intersect the molecular beam perpendicular to the direction of molecule motion. Laser-induced fluorescence is collected onto a photomultiplier tube (PMT) in photon counting mode. In the thermal source, the fluorescence is collimated, passed through an interference filter, and then spatially filtered and collected. In the cryogenic source, the fluorescence is transferred to the PMT by a light pipe, followed by an interference filter.

III Hyperfine and Rotation Hamiltonian

III.1 Quantum Numbers in the $X-B$ Spectrum

Thallium has two common isotopes, 203Tl (30 % natural abundance) and 205Tl (70 %), both with nuclear spin $I_{1}$ $=$ $1/2$ Lide (1997). Fluorine’s only isotope is 19F, also with nuclear spin $I_{2}$ $=$ $1/2$. We describe the $B$ state of TlF using the Hund’s case (c) basis and the coupling scheme:

[TABLE]

where $J$ is the total angular momentum of the molecule less nuclear spin. Hence each rotational state with quantum number $J$ has associated HF states with $F_{1}$ $=$ $J\pm\frac{1}{2}$ and $F$ $=$ $J-1$, $J$, $J$, and $J+1$.

The Hund’s case (c) basis kets $\ket{c}$ are:

[TABLE]

Here, $\Omega$ is the projection of $J$ on the internuclear axis, $m_{F}$ is the projection of $F$ in the lab frame, and $P=\pm 1$ is the state parity. Following the convention of Herzberg Herzberg (1950), we refer to states with $P$ $=$ $(-1)^{J}$ as $e$-parity and $P=(-1)^{J+1}$ as $f$-parity states. As described below, the large HF interactions in the $B$ state mix neighboring rotational levels. We use rotational quantum number $J$ to describe states in the basis of Eq. 2, and label energy eigenstates in the case of large mixing by $\tilde{J}$ (i.e. $\tilde{J}=J$ in the absence of HF mixing). We denote the rotational quantum number in the ground state by $J_{g}$.

III.2 Magnetic Hyperfine

The largest HF effect is expected to be due to the magnetic HF interaction, described by the Hamiltonian $H_{\rm{mhf}}$:

[TABLE]

where $I$ $=$ $I_{1}$ or $I_{2}$; $L$ is the total electron orbital angular momentum; and $S$ is the total electron spin. The lower-case subscript corresponds to coordinates in the molecule fixed-frame, with $\hat{z}$ along the internuclear axis. In the limit of negligible coupling to other electronic states via $H_{\rm{mhf}}$, we may write the effective Hamiltonian in the form

[TABLE]

Here, $h_{\Omega}$ $=$ $a\langle L_{z}\rangle+(b+c)\langle S_{z}\rangle$, where $\langle\rangle$ corresponds to the expectation value of the operators in the electronic state of interest. For $I$ $=$ $I_{1}$,

[TABLE]

Similarly for $I=I_{2}$,

[TABLE]

III.3 Nuclear Spin-Rotation

The effective nuclear spin-rotation Hamiltonian is of the form $H_{\rm{nsr}}$ $=$ $c_{I}(\boldsymbol{I}\cdot\boldsymbol{J})$, where again $I=I_{1}$ or $I_{2}$. This arises from two physical mechanisms. The first is the coupling of the rotational motion of the nuclei to the nuclear spin magnetic moments. For an electronic state which is not strongly perturbed by other nearby states, this contribution dominates, and $c_{I}$ is quite small (for example, $c_{I}$(Tl) = 126.03 kHz and $c_{I}$(F) = 17.89 kHz in the $X^{1}\Sigma$ state in TlF Hinds and Sandars (1980)). The second contribution arises from second-order rotational coupling to other electronic states Okabayashi et al. (2012). This contribution likely dominates in the $B$ state, where levels with ${}^{3}\Delta$ and ${}^{3}\Sigma$ character are predicted to lie near in energy Balasubramanian (1985). In such a situation, $c_{I}$ can be significantly larger than when the main contribution comes from the first mechanism; for example, in 195PtF, $c_{I}$(Pt) = 3.11 MHz Okabayashi et al. (2012).

The nuclear spin-rotation matrix elements can be written as follows. For $I$ $=$ $I_{1}$:

[TABLE]

For $I=I_{2}$:

[TABLE]

Because there are discrepancies in the literature about the explicit form of these HF matrix elements, we derive them in Appendix C.

III.4 Additional Terms

Equations 5 and 6 are only strictly valid for an isolated electronic level. Perturbations by a nearby level can lead to extra terms in the effective Hamiltonian with the $J$-dependence of centrifugal terms, modeled by substituting $h_{1}\rightarrow h_{1}+h_{1D}J(J+1)$ Okabayashi and Tanimoto (2003). However, the diagonal matrix elements of the centrifugal magnetic HF and nuclear spin-rotation Hamiltonians are identical, and the effects of off-diagonal elements are too small to be discernable in our data. We choose to constrain $h_{1D}$ $=$ [math], and treat the fit constants $c_{I}$(Tl) and $c_{I}$(F) as empirical combinations of the two effects.

Rotational energy is modeled using the standard effective Hamiltonian $H_{\rm{rot}}$ Herzberg (1950):

[TABLE]

with diagonal matrix elements

[TABLE]

The $B$ state HF/rotation structure is expected to depend on the isotopologue and the $e$/$f$ parity. We therefore fit the parity states and each Tl isotope separately. This is equivalent to the analysis of Ref. Okabayashi et al. (2012), which used the substitution $h_{1}$ $\rightarrow$ $h_{1}\pm h_{1q}$ and and $c_{I}$ $\rightarrow$ $c_{I}\pm c_{Iq}$, with the upper (lower) sign used for the $e$- ($f$-) parity states.

IV Observed Spectral Features

IV.1 Line Identification

The HF structure of the ground $X^{1}\Sigma^{+}$ state is unresolved ($\sim$ 100 kHz) in our optical spectra. In the excited $B^{3}\Pi_{1}$ state, we expect 8 well-split isotope/HF sublevels for each rotational level $\tilde{J}$ and parity $P$, corresponding to $F_{1}$ $=$ $\tilde{J}\pm\frac{1}{2}$, $F$ $=$ $F_{1}\pm\frac{1}{2}$, and the two Tl isotopes. We label the 7 splittings between the 8 lines as $a,b,\dots g$ as shown in Fig. 1. In most cases, we easily identify lines associated with the two isotopologues, since their intensities are proportional to the isotopic abundance: we associate splittings $a,b,c$ with 205Tl; $e,f,g$ with 203Tl; and $d$ with the gap between the isotopologues. For a given isotopologue, the largest splitting (associated with the Tl nuclear spin projection) is $b$ or $f$. The doublets separated by $a,c$ ($e,g$) then correspond to the 19F nuclear spin projection for 205TlF (203TlF). As described in Section IV.2 below, in $\tilde{J}=1$ only, the separation of the levels due to the Tl spin projection is larger than the isotope shift; we use the convention of labeling the splitting between the highest 205TlF HF level and the lowest 203TlF HF level as $d$, and thus $d<0$ for transitions to $\tilde{J}$ $=$ $1$.

The $R$ and $P$ branch transitions are spectroscopically isolated and hence relatively easy to identify for $\tilde{J}>1$. For these states, 8-line multiplets similar to that displayed in Fig. 1 are found approximately centered on the locations predicted from the Dunham coefficients of Tiemann Tiemann (1988). Our measured splittings for the various values of $e$-parity excited states are listed in Table 1.

Assignment of the $Q$ branch transitions is more challenging. Due to the near equality of the rotational constants of the $X$ and $B$ states, the $Q$ branch transitions are generally not clearly separated. All of the lines between $Q_{1}$ and $Q_{60}$ (approximately 480 individual HF transitions) are contained in a frequency range of about 21 GHz (see Appendix D). We initially identified many of the $Q$ branch multiplets for $Q_{11}$–$Q_{34}$ using rotational constants $B_{0}$, $D_{0}$, and $H_{0}$ from Tiemann Tiemann (1988), then fine-tuned these parameters to obtain good agreement with all the identifiable $Q$ branch lines ($Q_{11}$–$Q_{68}$). Because of their lower abundance and correspondingly smaller signals, no similar identification was possible for the majority of the 203TlF $Q$ branch lines. Assuming the same rotational constants (scaled for reduced mass and including an isotopic offset), we have made some tentative assignments of 203TlF $Q_{60}$–$Q_{71}$.

The HF splitting is larger than the separation of $Q$ branch transitions for $Q_{1}$–$Q_{10}$, and clear identification of the lines again becomes problematic. Ignoring the splittings which are obscured by overlapping $Q$ branch lines eliminates nearly all $Q_{1}$–$Q_{10}$ data (except $Q_{4}$ and $Q_{9}$). However, superposing the data taken in this spectral region from both the thermal beam and the cryogenic beam provides additional information and leads to the tentative assignments shown in Fig. 2a. The resulting splittings for $f$-parity excited states are listed in Table 2.

We now discuss a number of unusual features in the $B$ state HF structure. Many of these observations are attributed to an unusually large value of $h_{1}$(Tl) (found to be $\approx$ 29 GHz), which we believe may be the largest observed HF interaction in any diatomic molecule.

IV.2 Large Hyperfine Splitting for $\tilde{J}=1$

The large value of $h_{1}$(Tl) is best illustrated by the enormous $b$ splitting in $\tilde{J}=1$ (Fig. 2b). While the HF structure of $\tilde{J}=1$ covers $\sim\!13$ GHz, all higher $Q$ branch lines to $\tilde{J}\geq 2$ observable in the cryogenic source ($Q_{2}$ to roughly $Q_{8}$) are contained in only roughly 2 GHz. The large splitting in $\tilde{J}=1$ means two of the four $R_{0}$ lines are actually lower in frequency than $Q_{2}$.

IV.3 J-mixing and “Extra” Lines

In addition to the expected $P$, $Q$ and $R$ branch lines, we observe additional quartets of lines near several low $P$ and $R$ branch lines (Fig. 3). These are assigned as nominal $\Delta J$ $=$ $-2,+2$ transitions ($O$ and $S$ branch, respectively). The presence of these lines is an indication that the $B$ state rotational levels are strongly mixed by the magnetic HF interaction. If $|h_{1}$(Tl)$|$ $\gg|h_{1}$(F)$|$, as expected since HF structure is enhanced in heavier species, then this $J$-mixing is only significant between states with the same values of $F_{1}$ and $F$. The eigenstates can be written in the form

[TABLE]

Note that $\ket{\tilde{J}=1,F_{1}=1/2,F=0,1}$ is a special case which does not experience $J$-mixing. For all other excited states, $J$-mixing leads to electric dipole-allowed $O$ and $S$ branch transitions. Because the rotational constants of the $X$ and $B$ states are nearly identical, the $O_{J}$ lines appear very close to $P_{2J-1}$, and the $S_{J}$ lines appear very close to $R_{2J+2}$. We label the splittings as $h,i,j$ for the $O$-branch and $k,l,m$ for the $S$-branch (see Fig. 3).

The $P$ and $R$ branches target the excited state $e$-parity levels. We check our line assignments by comparing splittings in $R_{J}$ with those in $P_{J+2}$, which share a common excited state and thus should have the same observed HF splittings. The $O$, $Q$, and $S$ branches target the excited state $f$-parity levels. The $Q_{J}$ branch HF structure must be compared to both $O_{J+2}$ and $S_{J-2}$ branches, as these lines address only one value of $F_{1}$ in the excited state due to the selection rule $\Delta F_{1}$ $=$ $\pm 1,0$.

As an aside, we note that the 19F magnetic HF interaction mixes states with the same $F_{1}$, and thus it is in principle possible to drive electric dipole transitions with $\Delta J=-3,+3$, which we call the $N$ and $T$ branches, respectively. For nearly equal ground and excited state rotational constants, we expect two $T_{J}$ lines – one for each isotope, and split by $d+e+f+g$ for $R_{J+2}$ – to appear very close to $R_{3J+5}$ (and similarly, $N_{J}$ lines near $P_{3J-3}$). Because the majority of the molecules produced in our cryogenic buffer gas beam source are in $J_{g}=0,1$, we searched for $T_{0}$ and $T_{1}$ lines around $R_{5}$ and $R_{8}$, respectively. However, we failed to detect any such transitions.

IV.4 Inverted $F_{1}$ Doublet for $\tilde{J}\geq 2$

By examining the patterns in the data we can infer the ordering of the energy levels within different quartets of lines. The $k$ splitting of the $S$ branch lines is consistent with the $c$ splitting (upper fluorine doublet splitting) of the $Q$ branch, indicating that $\ket{\tilde{J},F_{1}=\tilde{J}-\frac{1}{2}}$ lies higher in energy than $\ket{\tilde{J},F_{1}=\tilde{J}+\frac{1}{2}}$ for all observed $S$ branch lines, corresponding to $\tilde{J}$ $=$ $2$–5. However, the $h$ splitting for $O_{3}$ is also consistent with the $c$ splitting of $Q_{1}$. Combined with the fact that $a>c$ for all lines connected to $\tilde{J}=1$, but $a<c$ for $\tilde{J}\geq 2$, we conclude that the ordering of $F_{1}$ levels is regular only in $\tilde{J}$ $=$ $1$, and inverted for $\tilde{J}$ $\geq$ 2. Also of note is that the $b$ splitting is $\sim\!29\times$ larger in $\tilde{J}$ $=$ $1$ than in $\tilde{J}$ $=$ $2$. Naively, the $J$-scaling of diagonal matrix elements of $H^{\rm{eff}}_{\rm{mhf}}$ (Eq. 5) would have led us to expect the $b$ splitting to only be $\sim 2\times$ larger in $\tilde{J}$ $=$ $1$ than in $\tilde{J}$ $=$ $2$.

These peculiar features can be explained by an exceptionally large value of $h_{1}$(Tl). Consider a simplified system where only $H_{\rm{mhf}}$(Tl) (Eq. 4) and $H_{\rm{rot}}$ $=$ $B\boldsymbol{J}^{2}$ are present. In the basis of states $\ket{J=F_{1}\pm 1/2,F_{1}}$ with $F_{1}\geq 3/2$, the interaction is characterized by the $2\times 2$ Hamiltonian

[TABLE]

By diagonalizing $H_{2\times 2}$, we find the energy $E_{F_{1}}^{\pm}$ of eigenstate $\ket{\tilde{J}=F_{1}\pm 1/2,F_{1}}$ is given by

[TABLE]

The energy difference between states $\ket{\tilde{J},F_{1}=\tilde{J}\pm 1/2}$ vanishes for $\tilde{J}\geq 2$ when $h$ $=$ 4$$B (see Fig. 4). The energy ordering of the $F_{1}$ doublets reverses for $h_{1}$ $>$ $4B$ for all $\tilde{J}\geq 2$, but not for $\tilde{J}=1$. For $h_{1}\simeq 4B$, the splitting of $\tilde{J}=1$ is much greater than that of all other states. These features are consistent with our observations and our fit value h_{1}\rm{(Tl)}$$/B_{0} $=$ $4.3$.

IV.5 Fitting

The best-fit values for the rotational and HF parameters are provided in Table 3. The $B$ state rotational parameters are obtained by subtracting off the precisely-known $X$ state rotational energy from the weighted center of the HF quartet for each isotope, then fitting to a polynomial in $\tilde{J}(\tilde{J}+1)$. For the $f$-parity, we perform a cubic fit to $Q_{11}$–$Q_{60}$. For the $e$-parity, we fit to $R_{1}$–$R_{8}$, and only the linear ($B_{0}$) term is statistically significant. The $R_{0}$ lines are excluded from the fit as they are observed to deviate strongly from the rotational progression as discussed above. The $e$-parity level $B_{0}$ constants are observed to match the expected scaling with molecular reduced mass $\mu$: ${}^{203}B_{0}/^{205}B_{0}$ $=$ $1.00080(4)$, while $\mu_{205}/\mu_{203}=1.00084$.

The HF fit parameters are obtained by diagonalizing the effective Hamiltonian $H$ $=$ H_{\rm{rot}}+$$H_{\rm{mhf}}(\rm{Tl})+$$H_{\rm{mhf}}(\rm{F})+$$H_{\rm{nsr}}(\rm{Tl})+$$H_{\rm{nsr}}(\rm{F}) for each value of $F$. The HF parameters for the $f$- ($e$)-parity states were fit to the observed $Q$ branch (average of the $P$ and $R$ branch) splitting, weighted by their assigned uncertainties. In all cases, we find $c_{I}$(F) to be consistent with zero within an uncertainty of 0.1 MHz.

The uncertainty in the splittings measured in the thermal beam is 3 MHz. For the cryogenic beam data, the Fabry-Perot ramp was less linear, and we assign an uncertainty of 8 MHz to these splittings. We fit to the thermal beam data if available, and the cryogenic beam data when not. In the instances where lines were measured in both setups, the agreement is typically within 10 MHz. As a check of our $Q$ branch splitting assignments where one or both lines are degenerate with other $Q$ branch lines (Table 2 bold data), a fit to the $f$-parity levels was performed excluding these data. However, this did not change the fit parameters within the assigned uncertainty. The excellent agreement between the data and the line centers predicted by the best-fit Hamiltonian (Fig. 2a) provides additional support for our tentative $Q$ branch line assignments.

There is fair agreement between the 205TlF and 203TlF isotopologues on the value of the 19F doublet splittings. The magnetic HF parameter $h_{1}$(Tl) should be proportional to the nuclear $g$-factor $g_{N}$, and we find excellent agreement for $e$-parity: $h_{1}(^{203}$Tl)/$h_{1}(^{205}$Tl) = 0.9904(2), while $g_{203}/g_{205}$ $=$ $0.990260$ Lide (1997). Hyperfine anomalies are known to occur at the $10^{-4}$ level in neutral Tl atoms Mårtensson-Pendrill (1995); Richardson et al. (2000), but is outside the precision of this study.

Finally, we extract the electronic isotope shift $\Delta\nu_{\rm{el}}(203-205)$ for the $X$ $\rightarrow$ $B$ transition Knöckel and Tiemann (1984). By averaging the shift for each HF transition for all $P$ and $R$ branch lines and subtracting off the calculated rovibrational contributions, we find $\Delta\nu_{\rm{el}}(203-205)$ $=$ $+3309(9)$ MHz. Because we only measure two Tl isotopes, we are unable to deconvolve the mass- and field-shift contributions to this quantity Knöckel and Tiemann (1984).

V Optical Cycling

V.1 Applications and Requirements

The ability to scatter many photons per molecule can be used for efficient detection and for application of optical forces. The number of optical cycles required before leaking into an uncoupled state depends on the application. For instance, simple state-selective detection of molecules in a beam often does not require a highly closed cycle. A typical laser-induced fluorescence setup can achieve overall detection efficiency for emitted photons (including detector quantum efficiency) of $\eta\sim 1$–10% Barry et al. (2014). Hence, to detect molecules with near-unit efficiency, we only require an optical cycling scheme closed to (very roughly) $1/\eta\!\sim\!10$–100 cycles. Applying significant optical forces requires a more closed cycle: $\sim\!10^{3}$ optical cycles are required to achieve transverse Doppler cooling of a molecular beam to milliKelvin temperatures Shuman et al. (2010), and $\sim\!10^{4}$ optical cycles are needed to slow a molecular beam to a near stop Barry et al. (2012) or to magneto-optically trap molecules Barry et al. (2014).

In this section, we consider vibrational and rotational branching effects which can limit the number of photons scattered on a given spectral line of the TlF $X$ $\rightarrow$ $B$ transition. We denote the vibrational branching fraction $v_{e}$ $\rightarrow$ $v_{g}$ by $b_{v_{e}v_{g}}$, and rotational branching fraction $\tilde{J}$ $\rightarrow$ $J_{g}$ (irrespective of vibrational quantum number) by $r_{\tilde{J}J_{g}}$. We then propose a few optical cycling schemes which can be used in TlF for specific applications.

V.2 Rotational Branching

As discussed in Section IV.3, HF interactions mix states of the same total angular momentum $F$. In Table 4, we present the eigenstates of $B(v_{e}$ = $0)$ in the basis of Eq. 2 calculated from the best-fit spectroscopic constants in Table 3. With the state admixture in hand, it is straightforward to calculate rotational branching fractions $r_{\tilde{J}J_{g}}$ using the line strengths $S$ Townes and Schawlow (1955) for each excited basis state.

V.3 Vibrational Branching

Previous measurements of vibrational branching in TlF found favorable branching fractions $b_{v_{e}v_{g}}$ for cycling from the $B(v_{e}=0$ state: $b_{00}$ = $0.99$, $b_{01}$ $<$ $0.0002$, and $b_{02}$ = $0.0110(6)$ Hunter et al. (2012). Measurements of branching fractions to $v_{g}$ $\geq$ 3 were limited by experimental sensitivity and the availability of narrow bandpass interference filters at the needed wavelengths. Predicted values for these branching fractions using Morse and Rydberg-Klein-Rees (RKR) potentials, given in Table 5, were uncertain at the level required to determine which of these transitions could be neglected during longitudinal cooling. Here we present precision measurements of $b_{03}$, $b_{04}$, $b_{05}$, and $b_{06}$.

Measurements of the branching fractions are done using a modified version of the thermal beam in order to take advantage of the better molecular beam intensity stability. Pulsed excitation is used, and fluorescence is only detected after the laser pulse, eliminating scattered-light backgrounds. The exciting laser light is produced by a pulsed dye laser with Coumarin 540A dye pumped by a 355 nm Nd:YAG laser with pulse duration of 10 ns and 10 Hz repetition rate. The output of the dye laser is frequency doubled to produce 271.7 nm light. Measurements are made with the broadband pulsed system tuned to the largest fluorescence signal that occurs, near a large pile-up of rotational transitions between $Q_{30}$ and $Q_{52}$ with a bandhead that occurs at about $Q_{43}$. Background measurements are made by tuning to the high frequency side of this bandhead where virtually no fluorescence occurs.

Bandpass interference filters centered at the wavelengths corresponding to the $B(0)$ $\rightarrow$ $X(0$–6$)$ transitions are placed before the “signal” PMT (above the vacuum chamber) to isolate fluorescence from the different vibrational bands. A second “normalization” PMT (below the chamber) contains only 271.7 nm interference filters and is used to measure fluorescence from the main transition ($B(0)$ $\rightarrow$ $X(0)$) at all times. This allows us to eliminate fluctuations in laser and molecular beam intensity from our measurements. To avoid saturation problems with the photon counting, we limit the average number of photons collected by either PMT to be $<1$ per laser pulse.

To measure a particular branching ratio we alternate between the corresponding filters, and count photons after subtracting off background fluorescence, normalizing with respect to the calibration signal, and taking into account the filter transmissions. Measurement of the ratio $b_{02}/b_{00}$ is accomplished with a modest molecular beam flux by dividing the fluorescence signals when alternately filtering for $B(0)$ $\rightarrow$ $X(2)$ and $B(0)$ $\rightarrow$ $X(0)$. However, the small rate of the decay to $v_{g}\neq 0,2$ requires higher molecular beam fluxes to achieve adequate statistical precision. To avoid saturating the normalization PMT from these larger fluorescence signals, additional attenuating filter are added. To avoid saturating the signal detector at these higher fluxes, all branching fractions $b_{0v}$ for $v_{g}\neq 0,2$ are measured by comparing the normalized signal for $B(0)$ $\rightarrow$ $X(v_{g})$ to that of $B(0)$ $\rightarrow$ $X(2)$. Their measured ratio is then multiplied by $b_{02}$ as measured above to determine $b_{0v_{g}}$. The attenuation on the normalization transition is chosen so that the statistical significance of the measurement remains dominated by the number of photons counted in the signal detector.

We calculate the experimental branching fractions displayed in Table 5 using the assumption that the measured branching fractions account for all of the significant vibrational branches from $v_{e}$ $=$ [math]. It is worth noting that the uncertainty associated with $b_{02}$ is larger than that quoted in our earlier publication Hunter et al. (2012). The increased uncertainty is due to the identification of a systematic error associated with a changing amount of light reflected into the normalization PMT when the interference filter in the signal PMT is switched between monitoring the $v_{g}=0$ and $v_{g}=2$ transitions. This effect was likely also present in our earlier measurement.

V.4 Optical Cycling Schemes

We now examine a few specific, useful examples of optical cycling $\ket{X^{1}\Sigma^{+},J_{g}}$ $\rightarrow$ $\ket{B^{3}\Pi_{1},\tilde{J},F_{1},F,P=f}$ (which we abbreviate below as $\ket{J_{g}}$ $\rightarrow$ $\ket{\tilde{J},F_{1},F}$). First consider $\ket{1}$ $\rightarrow$ $\ket{1,1/2,0}$ (Fig. 5a). As there are no other $F$ $=$ [math] states, the upper state of this transition is unmixed and all quantum numbers are exact to a very high extent (Table 4). Electric dipole and parity selection rules then dictate that the excited state can only decay to $J_{g}=1$.

Now consider $\ket{1}$ $\rightarrow$ $\ket{1,1/2,1}$ (Fig. 5b). This transition may be of interest in applications where a high photon scattering rate is desirable, as the higher excited state degeneracy $2F+1$ is expected to allow for roughly $3\times$ higher photon scattering rate than the case $\ket{1}$ $\rightarrow$ $\ket{1,1/2,0}$ Norrgard et al. (2016); Norrgard (2016). However, the $\ket{1,1/2,1}$ excited state has $|0.0180|^{2}\simeq 3\times 10^{-4}$ fractional $J=2$ character (Table 4). Decays from $J$ $=$ $2$ go to the desired $J_{g}=1$ (2/5 of the time) or to $J_{g}=3$ (3/5 of the time). Hence, only $1/r_{13}\simeq 5000$ photons can be scattered before molecules are lost to the uncoupled $J_{g}=3$ state. This loss is inconsequential for molecule detection applications, but is unacceptable for laser slowing and cooling. This population could be recovered by repumping $J_{g}=3$ with an additional laser tuned to the $\ket{3}$ $\rightarrow$ $\ket{2,3/2,1,e}$ transition. In this case, all electric dipole decay paths to $v_{g}$ = 0 would again be optically coupled.

Several possible optical cycling schemes exist which are closed to $\sim\!10^{4}$ photon scatters, sufficient for laser slowing or trapping. We present one such scheme in Fig. 5c. The main cycling transition is $\ket{1}$ $\rightarrow$ $\ket{1,1/2,0}$ to minimize rotational branching. Lasers $\mathcal{L}_{v_{g}v_{e}}$ (corresponding to the transition $X(v_{g})$ $\rightarrow$ $B(v_{e})$) may be added to the optical cycle in order of decreasing importance ($\mathcal{L}_{00}$, $\mathcal{L}_{20}$, $\mathcal{L}_{40}$, $\mathcal{L}_{10}$), until the system is sufficiently closed for the intended application. The resolved excited state HF structure allows the three strongest off-diagonal vibronic decays to be repumped through $v_{e}\!=\!0$ to take advantage of the near-unity $b_{00}$ branching fraction. We calculate that with only the $\mathcal{L}_{00}$ laser, $\sim$ 90 optical cycles may be achieved before molecules decay into a higher vibrational level. This should be sufficient for high-efficiency detection. The addition of one repump laser $\mathcal{L}_{20}$ should be sufficient to achieve transverse cooling (closed to $\sim 3600$ cycles). All four lasers shown in Fig. 5c would be necessary to achieve laser slowing or trapping. In this scheme, we expect to scatter $\sim\!10^{4}$ photons before decaying to unpumped levels.

VI Conclusions

In addition to spectroscopically identifying the $Q_{1}$ cycling transition of TlF, we have characterized the HF structure of the $B^{3}\Pi_{1}$ state. Of particular note is the large magnetic HF interaction of the Tl nuclear spin. With the exception of the lowest rotational and HF sublevel, the HF interactions significantly mix neighboring rotational levels of the $B$ state, and lead to additional rotational branching.

TlF appears to be an excellent molecule for optical cycling. In particular, rotational and vibrational branching fractions presented here indicate $\sim\!90$ optical cycles may be achieved using a single laser. Efficient detection with a single laser should be possible in a symmetry violation measurement using TlF. In addition, laser cooling and trapping should be feasible with four or fewer lasers.

While the measurements presented here show that the $X$ $\rightarrow$ $B$ system of TlF should allow for highly closed optical cycling, a number of considerations should be accounted for when choosing an appropriate cycling scheme in order to achieve a high photon scattering rate. We plan to detail these considerations in a future paper.

Financial support was provided by the Army Research Office, the John Templeton Foundation, the Heising-Simons Foundation, and by NSF Grants No. PHY1205824 and PHY1519265. The authors thank M. Kozlov and T. Steimle for helpful discussions. S.S.A. and N.W. would like to thank Amherst College for Summer Research Awards.

Appendices

VI.1 Vibrational Branching Further Details

In measuring the branching ratios in Section V.3, we use a number of interference filters to isolate fluorescence from individual vibronic bands. Table 6 provides further details on the filters used to monitor each transition.

VI.2 Racah Algebra Identities

Here we provide a number of useful identities and their equation numbers in Ref. Brown and Carrington (2003). In the following, $j_{1}$ and $j_{2}$ are general angular momenta which couple to form angular momentum $j_{12}$. Rank-$k$ tensor operators $T^{k}(\boldsymbol{A_{1}})$ and $T^{k}(\boldsymbol{A_{2}})$ act on $j_{1}$ and $j_{2}$, respectively. Tensors are represented in the spherical basis, with index $p$ ($q$) for the lab frame (molecule-fixed frame) coordinate system.

Transformation from molecule-fixed frame to lab frame:

[TABLE]

Scalar product of two tensor operators:

[TABLE]

Spectator Theorem:

[TABLE]

Reduced matrix element of a rank-1 tensor operator:

[TABLE]

Reduced matrix element of the partially reduced Wigner rotation matrix:

[TABLE]

VI.3 Evaluating Matrix Elements

Here we derive the HF matrix elements given in Section III. We have found references Brown and Carrington (2003); Okabayashi and Tanimoto (2003); Okabayashi et al. (2012); Knurr et al. (2009) helpful in understanding which terms are likely important in the HF structure Hamiltonian. Because Ref. Brown and Carrington (2003) contains all necessary fomulae for Racah algebra for diatomic molecules, we use the short-hand $\underset{(x.xx)}{=}$ for “is equivalent by Ref. Brown and Carrington (2003) Eq. (x.xx)”.

We begin by rotating the operator of Equation 4 into the lab frame:

[TABLE]

For $I=I_{1}$,

[TABLE]

which matches Ref. Okabayashi et al. (2012) Eq. (6). Note that when employing (5.173) to get the first line of Eq. 15, both the Wigner rotation matrix $\mathscr{D}$ and $T^{1}(\boldsymbol{I_{1}})$ are being reduced in the lab frame (indexed by $p$). Specifically, the term $\matrixelement{J,\Omega}{|\mathscr{D}^{(1)}_{\cdot,0}(\omega)^{*}|}{J^{\prime},\Omega^{\prime}}$ is the partially reduced Wigner rotation matrix (reduced in the lab frame, but not the molecule-fixed frame), whose value is given by Ref. Brown and Carrington (2003) Eq. 5.186. For $I=I_{2}$, Eq. 14 yields

[TABLE]

which again matches Ref. Okabayashi et al. (2012) Eq. (6).

The nuclear spin-rotation matrix elements can be written as follows. For $I=I_{1}$:

[TABLE]

Now for $I=I_{2}$:

[TABLE]

Results 17 and 18 match Ref. Brown and Carrington (2003) (8.290) and (8.291), respectively, as well as Ref. Knurr et al. (2009) (5). However, Eq. 17 differs from Ref. Okabayashi et al. (2012) by a factor of $[J(J+1)(2J+1)]^{1/2}$.

VI.4 Full $Q$ Branch Spectrum

In Fig. 6, we include the full $Q$ branch spectrum of the $X$ $\rightarrow$ $B$ transition in TlF. Line centers predicted by the best-fit rotational constants in Table 3 are in excellent agreement with the observed spectrum.